Non-Noether symmetries in Hamiltonian Dynamical Systems
George Chavchanidze
Department of Theoretical Physics,
A. Razmadze Institute of Mathematics,
1 Aleksidze Street, Tbilisi 0193, Georgia
Abstract.
We discuss geometric properties of non-Noether symmetries and
their possible applications in integrable Hamiltonian systems.
Correspondence between non-Noether symmetries and conservation laws
is revisited. It is shown that in regular Hamiltonian systems
such symmetries canonically lead to Lax pairs on the algebra
of linear operators on cotangent bundle over the phase space.
Relationship between non-Noether symmetries and other widespread geometric
methods of generating conservation laws such as bi-Hamiltonian formalism,
bidifferential calculi and Frölicher-Nijenhuis geometry is considered.
It is proved that the integrals of motion associated with a
continuous non-Noether symmetry are in involution whenever the
generator of the symmetry satisfies a certain Yang-Baxter type equation.
Action of one-parameter group of symmetry on algebra of integrals of motion
is studied and involutivity of group orbits is discussed.
Hidden non-Noether symmetries of Toda chain, Korteweg-de Vries equation,
Benney system, nonlinear water wave equations and Broer-Kaup system
are revealed and discussed.
Keywords: Non-Noether symmetry; Conservation law; bi-Hamiltonian system; Bidifferential calculus; Lax pair; Frölicher-Nijenhuis
operator; Korteweg-de Vries equation; Broer-Kaup system; Benney system; Toda chain
MSC 2000: 70H33; 70H06; 58J70; 53Z05; 35A30
1. Introduction
Symmetries play essential role in dynamical systems, because they usually simplify
analysis of evolution equations and often provide quite elegant solution of problems that otherwise would
be difficult to handle. In Lagrangian and Hamiltonian dynamical systems special role is played
by Noether symmetries — an important class of symmetries that leave action invariant
and have some exceptional features. In particular, Noether symmetries deserved
special attention due to celebrated Noether's theorem, that established correspondence
between symmetries, that leave action functional invariant, and conservation laws
of Euler-Lagrange equations. This correspondence can be extended to Hamiltonian
systems where it becomes more tight and evident then in Lagrangian case and gives rise
to Lie algebra homomorphism between Lie algebra of Noether symmetries and algebra of
conservation laws (that form Lie algebra under Poisson bracket).
Role of symmetries that are not of Noether type has been suppressed for quite a long time.
However, after some publications of Hojman, Harleston, Lutzky and others
(see [16], [36], [39],
[40], [49]-[57])
it became clear that non-Noether symmetries also can play important role in
Lagrangian and Hamiltonian dynamics. In particular, according to Lutzky
[51], in Lagrangian dynamics there is definite correspondence between non-Noether symmetries and
conservation laws. Moreover, each generator of non-Noether symmetry
may produce whole family of conservation laws (maximal number of conservation laws that can
be associated with non-Noether symmetry via Lutzky's theorem is equal to the dimension of
configuration space of Lagrangian system). This fact makes non-Noether symmetries especially
valuable in infinite dimensional dynamical systems, where potentially one can recover
infinite sequence of conservation laws knowing single generator of non-Noether symmetry.
Existence of correspondence between non-Noether symmetries and conserved quantities
raised many questions concerning relationship among this type of symmetries and
other geometric structures emerging in theory of integrable models.
In particular one could notice suspicious similarity between the method of constructing
conservation laws from generator of non-Noether symmetry and
the way conserved quantities are produced in either Lax theory, bi-Hamiltonian formalism,
bicomplex approach or Lenard scheme.
It also raised natural question whether set of conservation laws associated with non-Noether
symmetry is involutive or not, and since it appeared that in general it may not be involutive,
there emerged the need of involutivity criteria, similar to Yang-Baxter equation used in Lax theory
or compatibility condition in bi-Hamiltonian formalism and bicomplex approach.
It was also unclear how to construct conservation laws in case of infinite dimensional
dynamical systems where volume forms used in Lutzky's construction are no longer well defined.
Some of these questions were addressed in papers [11]-[14],
while in the present review we would like to summarize all these issues and to provide some
examples of integrable models that possess non-Noether symmetries.
Review is organized as follows. In first section we briefly recall some aspects of geometric
formulation of Hamiltonian dynamics. Further, in second section, correspondence
between non-Noether symmetries and integrals of motion in regular Hamiltonian systems is
discussed. Lutzky's theorem is reformulated in terms of bivector fields
and alternative derivation of conserved quantities suitable for computations in infinite
dimensional Hamiltonian dynamical systems is suggested. Non-Noether symmetries of
two and three particle Toda chains are used to illustrate general theory.
In the subsequent section geometric formulation of Hojman's theorem [36]
is revisited and some examples are provided. Section 4 reveals correspondence between
non-Noether symmetries and Lax pairs. It is shown that non-Noether symmetry canonically
gives rise to a Lax pair of certain type. Lax pair is explicitly constructed in terms
of Poisson bivector field and generator of symmetry. Examples of Toda chains are discussed.
Next section deals with integrability issues. An analogue of Yang-Baxter equation
that, being satisfied by generator of symmetry, ensures involutivity of set
of conservation laws produced by this symmetry, is introduced.
Relationship between non-Noether symmetries and bi-Hamiltonian systems
is considered in section 6. It is proved that under certain conditions,
non-Noether symmetry endows phase space of regular Hamiltonian system with
bi-Hamiltonian structure. We also discuss conditions under which non-Noether
symmetry can be "recovered" from bi-Hamiltonian structure.
Theory is illustrated by example of Toda chains. Next section is devoted to
bicomplexes and their relationship with non-Noether symmetries. Special kind
of deformation of De Rham complex induced by symmetry is constructed in terms of
Poisson bivector field and generator of symmetry.
Samples of two and three particle Toda chain are discussed.
Section 8 deals with Frölicher-Nijenhuis recursion operators.
It is shown that under certain condition non-Noether symmetry
gives rise to invariant Frölicher-Nijenhuis operator on tangent
bundle over phase space.
The last section of theoretical part contains some remarks on action of one-parameter
group of symmetry on algebra of integrals of motion. Special attention is devoted to
involutivity of group orbits.
Subsequent sections of present review provide examples of integrable models
that possess interesting non-Noether symmetries. In particular section 10 reveals
non-Noether symmetry of n-particle Toda chain. Bi-Hamiltonian structure,
conservation laws, bicomplex, Lax pair and Frölicher-Nijenhuis recursion
operator of Toda hierarchy are constructed using this symmetry.
Further we focus on infinite dimensional integrable Hamiltonian systems emerging
in mathematical physics. In section 11 case of Korteweg-de Vries
equation is discussed. Symmetry of this equation is identified and used in construction
of infinite sequence of conservation laws and bi-Hamiltonian structure of
KdV hierarchy. Next section
is devoted to non-Noether symmetries of integrable systems of nonlinear water wave equations,
such as dispersive water wave system, Broer-Kaup system and dispersiveless long wave system.
Last section focuses on Benney system and its non-Noether symmetry, that appears to be local,
gives rise to infinite sequence of conserved densities of Benney hierarchy and endows it with
bi-Hamiltonian structure.
2. Regular Hamiltonian systems
The basic concept in geometric formulation of Hamiltonian dynamics
is notion of symplectic manifold. Such a manifold plays the role of
the phase space of the dynamical system and therefore many properties
of the dynamical system can be quite effectively investigated in the framework
of symplectic geometry. Before we consider symmetries of the Hamiltonian dynamical
systems, let us briefly recall some basic notions from symplectic geometry.
The symplectic manifold is a pair (M, ω)
where M is smooth even dimensional manifold and ω
is a closed
dω = 0(1)
and nondegenerate 2-form on M. Being nondegenerate means that
contraction of arbitrary non-zero vector field with ω does not vanish
iXω = 0 ⇔ X = 0(2)
(here iX denotes contraction of the vector field X
with differential form). Otherwise one can say that ω
is nondegenerate if its n-th outer power does not vanish
(ωn ≠ 0) anywhere on M.
In Hamiltonian dynamics M is usually phase space of classical dynamical system
with finite numbers of degrees of freedom and the symplectic form ω
is basic object that defines Poisson bracket structure, algebra of Hamiltonian vector fields
and the form of Hamilton's equations.
The symplectic form ω naturally defines isomorphism between vector fields
and differential 1-forms on M (in other words tangent bundle TM
of symplectic manifold can be quite naturally identified with
cotangent bundle T*M).
The isomorphic map Φω from TM into
T*M is obtained by taking contraction
of the vector field with ω
Φω: X → − iXω
(3)
(minus sign is the matter of convention). This isomorphism gives rise to natural classification
of vector fields. Namely, vector field Xh is said to be Hamiltonian
if its image is exact 1-form or in other words if it satisfies Hamilton's equation
iXhω + dh = 0
(4)
for some function h on M.
Similarly, vector field X is called locally Hamiltonian if it's image is closed 1-form
iXω + u = 0, du = 0
(5)
One of the nice features of locally Hamiltonian vector fields, known as Liouville's theorem,
is that these vector fields preserve symplectic form ω.
In other words Lie derivative of the symplectic form ω
along arbitrary locally Hamiltonian vector field vanishes
LXω = 0 ⇔ iXω + du = 0, du = 0
(6)
Indeed, using Cartan's formula that expresses Lie derivative in terms of contraction and
exterior derivative
LX = iXd + diX
(7)
one gets
LXω = iXdω + diXω =
diXω
(8)
(since dω = 0) but according to the definition of locally Hamiltonian
vector field
diXω = − du = 0
(9)
So locally Hamiltonian vector fields preserve ω and vise versa,
if vector field preserves symplectic form ω then it is locally Hamiltonian.
Clearly, Hamiltonian vector fields constitute subset of locally Hamiltonian ones since
every exact 1-form is also closed. Moreover one can notice that Hamiltonian vector fields form
ideal in algebra of locally Hamiltonian vector fields. This fact can be observed as follows.
First of all for arbitrary couple of locally Hamiltonian vector fields X, Y
we have LXω = LYω = 0 and
LXLYω − LYLXω
= L[X , Y]ω = 0
(10)
so locally Hamiltonian vector fields form Lie algebra (corresponding Lie bracket is ordinary
commutator of vector fields). Further it is clear that for arbitrary Hamiltonian vector field
Xh and locally Hamiltonian one Z one has
LZω = 0
(11)
and
iXhω + dh = 0
(12)
that implies
LZ(iXhω + dh)
= L[Z , Xh]ω + iXhLZω +
dLZh
= L[Z , Xh]ω + dLZh = 0
(13)
thus commutator [Z , Xh] is Hamiltonian vector field
XLZh,
or in other words Hamiltonian vector fields form ideal in algebra of locally
Hamiltonian vector fields.
Isomorphism Φω can be extended to
higher order vector fields and differential forms by linearity and multiplicativity.
Namely,
Φω(X ∧ Y) =
Φω(X) ∧ Φω(Y)
(14)
Since Φω is isomorphism, the symplectic form ω
has unique counter image W known as Poisson bivector field.
Property dω = 0 together with non degeneracy implies that bivector
field W is also nondegenerate (Wn ≠ 0) and satisfies
condition
[W , W] = 0
(15)
where bracket [ , ] known as Schouten bracket or supercommutator, is actually
graded extension of ordinary commutator of vector fields to the case of multivector fields,
and can be defined by linearity and derivation property
[C1 ∧ C2 ∧ ... ∧ Cn ,
S1 ∧ S2 ∧ ... ∧ Sn] =
(− 1)p + q[Cp , Sq] ∧
C1 ∧ C2 ∧ ... ∧ Ĉp ∧ ... ∧ Cn
∧ S1 ∧ S2 ∧ ... ∧ Ŝq ∧ ...∧ Sn
(16)
where over hat denotes omission of corresponding vector field.
In terms of the bivector field W Liouville's theorem mentioned above can be
rewritten as follows
[W(u) , W] = 0 ⇔ du = 0
(17)
for each 1-form u. It follows from graded Jacoby identity satisfied by Schouten
bracket and property [W , W] = 0 satisfied by Poisson bivector field.
Being counter image of symplectic form, W gives rise to map
ΦW, transforming differential 1-forms into vector fields,
which is inverted to the map Φω and is defined by
ΦW: u → W(u); ΦWΦω = id
(18)
Further we will often use these maps.
In Hamiltonian dynamical systems Poisson bivector field is geometric object that
underlies definition of Poisson bracket — kind of Lie bracket on algebra of
smooth real functions on phase space. In terms of bivector field W
Poisson bracket is defined by
{f , g} = W(df ∧ dg)
(19)
The condition [W , W] = 0 satisfied by bivector field ensures that
for every triple (f, g, h) of smooth
functions on the phase space the Jacobi identity
{f{g , h}} + {h{f , g}} + {g{h , f}} = 0.
(20)
is satisfied.
Interesting property of the Poisson bracket is that map from algebra of real smooth functions
on phase space into algebra of Hamiltonian vector fields, defined by Poisson bivector field
f → Xf = W(df)
(21)
appears to be homomorphism of Lie algebras. In other words commutator of two vector fields
associated with two arbitrary functions reproduces vector field associated with Poisson
bracket of these functions
[Xf , Xg] = X{f , g}
(22)
This property is consequence of the Liouville theorem and definition of Poisson bracket.
Further we also need another useful property of Hamiltonian vector fields and Poisson bracket
{f , g} = W(df ∧ dg) = ω(Xf ∧ Xg) =
LXfg = − LXgg
(23)
it also follows from Liouville theorem
and definition of Hamiltonian vector fields and Poisson brackets.
To define dynamics on M one has to specify time evolution of observables
(smooth functions on M). In Hamiltonian dynamical systems time evolution
is governed by Hamilton's equation
ddtf = {h , f}
(24)
where h is some fixed smooth function on the phase space called Hamiltonian.
In local coordinate frame zb bivector field W
has the form
W = Wbc ∂∂zb ∧ ∂∂zc
(25)
and the Hamilton's equation rewritten in terms of local coordinates takes the form
żb = Wbc∂h∂zb
(26)
3. Non-Noether symmetries
Now let us focus on symmetries of Hamilton's equation (24).
Generally speaking, symmetries play very important role in Hamiltonian dynamics
due to different reasons. They not only give rise to conservation laws but
also often provide very effective solutions to problems that otherwise would be difficult
to solve. Here we consider special class of symmetries of Hamilton's equation
called non-Noether symmetries. Such a symmetries appear to be closely related to
many geometric concepts used in Hamiltonian dynamics including bi-Hamiltonian structures,
Frölicher-Nijenhuis operators, Lax pairs and bicomplexes.
Before we proceed
let us recall that each vector field E on the phase space generates
the one-parameter continuous group of transformations
gz = ezLE
(27)
(here L denotes Lie derivative)
that acts on the observables as follows
gz(f) = ezLE(f) =
f + zLEf + ½(zLE)2f + ⋯
(28)
Such a group of transformation is called symmetry of Hamilton's equation (24)
if it commutes with time evolution operator
ddt gz(f)
= gz(ddtf)
(29)
in terms of the vector fields this condition means that the generator
E of the group gz commutes with the vector field
W(h) = {h , }, i. e.
[E , W(h)] = 0.
(30) However we would like to consider more general
case where E is time dependent vector field on phase space. In this case
(30) should be replaced with
∂∂tE = [E , W(h)].
(31)
Further one should distinguish between groups of symmetry transformations generated by Hamiltonian,
locally Hamiltonian and non-Hamiltonian vector fields. First kind of symmetries
are known as Noether symmetries and are widely used in Hamiltonian dynamics due to their
tight connection with conservation laws. Second group of symmetries is rarely used.
While third group of symmetries that further will be referred
as non-Noether symmetries seems to play important role in integrability issues due to
their remarkable relationship with bi-Hamiltonian structures and
Frölicher-Nijenhuis operators. Thus if in addition to (30) the
vector field E does not preserve Poisson bivector field [E , W] ≠ 0 then
gz is called non-Noether symmetry.
Now let us focus on non-Noether symmetries. We would like to show that the presence of
such a symmetry essentially enriches the geometry of the phase space
and under the certain conditions can ensure integrability of the dynamical system.
Before we proceed let us recall that the non-Noether symmetry leads to a number of
integrals of motion. More precisely the
relationship between non-Noether symmetries and the conservation laws is described by
the following theorem. This theorem was proposed by Lutzky in [51].
Here it is reformulated in terms of Poisson bivector field.
Theorem 1.
Let (M , h) be regular Hamiltonian system on the 2n-dimensional
Poisson manifold M. Then, if the vector field E generates
non-Noether symmetry, the functions
Y(k) = Ŵk ∧ Wn − kWn k = 1,2, ... n
(32)
where Ŵ = [E , W], are integrals of motion.
Proof.
By the definition
Ŵk ∧ Wn − k = Y(k)Wn.
(33)
(definition is correct since the space of 2n degree multivector fields on 2n
degree manifold is one dimensional).
Let us take time derivative of this expression along the vector field W(h),
ddtŴk ∧ Wn − k =
(ddtY(k))Wn
+ Y(k)[W(h) , Wn]
(34)
or
k(ddtŴ) ∧ Ŵk − 1 ∧ Wn − k
+ (n − k)[W(h) , W] ∧ Ŵk ∧ Wn − k − 1
= (ddtY(k))Wn
+ nY(k)[W(h) , W] ∧ Wn − 1
(35)
but according to the Liouville theorem the Hamiltonian vector field preserves W i. e.
ddtW = [W(h) , W] = 0
(36)
hence, by taking into account that
ddtE= ∂∂tE + [W(h) , E] = 0
(37) we get
ddtŴ
=
ddt[E , W] = [ddtE , W] + [E[W(h) , W]] = 0.
(38)
and as a result (35) yields
ddtY(k)Wn = 0
(39)
but since the dynamical system is regular (Wn ≠ 0)
we obtain that the functions Y(k) are integrals of motion.
Remark 1. Instead of conserved quantities
Y(1) ... Y(n), the
solutions c1 ... cn of the secular equation
(Ŵ − cW)n = 0
(40)
can be associated with the generator of symmetry.
By expanding expression (40) it is easy to verify that the conservation laws
Y(k) can be expressed in terms of the integrals of motion
c1 ... cn in the following way
Y(k) =
(n − k)! k!n! ∑ms > mt cm1cm2 ⋯ cmk
(41)
Note also that conservation laws Y(k) can be also defined by means of
symplectic form ω using the following formula
Y(k) = (LEω)k ∧ ωn − kωn k = 1,2, ... n
(42)
Conservation laws c1 ... cn can be also derived from
the secular equation
(LEω − cω)n = 0
(43)
However all these expressions fail in case of infinite dimensional Hamiltonian systems
where the volume form
Ω = ωn
(44)
does not exist since n = ∞. But fortunately in these case one can define conservation laws using
alternative formula
C(k) = iWk(LEω)k
(45)
as far as it involves only finite degree differential forms
(LEω)k and well defined multivector fields
Wk.
Note that in finite dimensional case the sequence of conservation laws C(k)
is related to families of conservation laws Y(k) and ck in the
following way
C(k) =
(n − k)! k!n! ∑ms > mt cm1cm2 ⋯ cmk
= n!(n − k)! k! Y(k)
(46)
Note also that by taking Lie derivative of known conservation along the generator of
symmetry E one can construct new conservation laws
ddtY = LXhY = 0 ⇒
ddtLEY = LXhLEY =
LELXhY = 0
(47)
since [E , Xh] = 0.
Remark 2. Besides continuous non-Noether symmetries generated by non-Hamiltonian
vector fields one may encounter discrete non-Noether symmetries — noncannonical
transformations that doesn't necessarily form group but commute with evolution operator
ddt g(f) = g(ddtf)
(48)
Such a symmetries give rise to the same conservation laws
Y(k) = g(W)k ∧ Wn − kWn k = 1,2, ... n
(49)
Example 1.
Let M be R4 with coordinates
z1, z2, z3, z4 and Poisson bivector field
W =
∂∂z1 ∧ ∂∂z3 +
∂∂z2 ∧ ∂∂z4
(50)
and let's take the following Hamiltonian
h =
12z12 +
12z22 + ez3 − z4
(51)
This is so called two particle non periodic Toda model.
One can check that the vector field defined as
E = 4∑s = 1 Es∂∂zs
(52)
with components
E1 =
12z12 − ez3 − z4 −
t2(z1 + z2)ez3 − z4
E2 =
12z22 + 2ez3 − z4 +
t2(z1 + z2)ez3 − z4
E3 =
2z1 +
12z2 +
t2(z12 + ez3 − z4)
E4 = z2 − 12z1 +
t2(z22 + ez3 − z4)
(53)
satisfies (31) condition and as a result generates symmetry of the dynamical system.
The symmetry appears to be non-Noether with Schouten bracket [E , W] equal to
Ŵ = [E , W] = z1∂∂z1 ∧ ∂∂z3
+ z2∂∂z2 ∧ ∂∂z4
+ ez3 − z4 ∂∂z1 ∧ ∂∂z2 +
∂∂z3 ∧ ∂∂z4
(54)
Calculation of volume vector fields
Ŵk ∧ Wn − k gives rise to
W ∧ W = − 2 ∂∂z1 ∧ ∂∂z2 ∧ ∂∂z3 ∧ ∂∂z4
Ŵ ∧ W = − (z1 + z2)
∂∂z1 ∧ ∂∂z2 ∧ ∂∂z3 ∧ ∂∂z4
Ŵ ∧ Ŵ =
− 2(z1z2 − ez3 − z4)
∂∂z1 ∧ ∂∂z2 ∧ ∂∂z3 ∧ ∂∂z4
(55)
and the conservation laws associated with this symmetry are just
Y(1) = Ŵ ∧ WW ∧ W =
12(z1 + z2)
Y(2) = Ŵ ∧ ŴW ∧ W =
z1z2 − ez3 − z4
(56)
It is remarkable that the same symmetry is also present in higher dimensions.
For example in case where M is R6 with coordinates
z1, z2, z3, z4, z5, z6
(57)
Poisson bivector equal to
W = ∂∂z1 ∧ ∂∂z4 + ∂∂z2 ∧ ∂∂z5 +
∂∂z3 ∧ ∂∂z6
(58)
and the following Hamiltonian
h =
12z12 +
12z22 +
12z32 +
ez4 − z5 +
ez5 − z6
(59)
we still can construct symmetry similar to (53).
More precisely the vector field defined for arbitrary function F as
E = 6∑s = 1 Es∂∂zs
(60)
with components specified as follows
E1 =
12z12 − 2ez4 − z5 −
t2(z1 + z2)ez4 − z5
E2 =
12z22 + 3ez4 − z5 −
ez5 − z6 +
t2(z1 + z2)ez4 − z5
E3 =
12z32 + 2ez5 − z6 +
t2(z2 + z3)ez5 − z6
(61)
E4 =
3z1 + 12z2 + 12z3 +
t2(z12 + ez4 − z5)
E5 =
2z2 − 12z1 + 12z3 +
t2(z22 + ez4 − z5 +
ez5 − z6)
E6 =
z3 − 12z1 − 12z2 +
t2(z32 + ez5 − z6)
(62)
satisfies (31) condition and generates non-Noether symmetry of the dynamical system
(three particle non periodic Toda chain).
Calculation of Schouten bracket [E , W] gives rise to expression
Ŵ = [E , W] = z1∂∂z1 ∧ ∂∂z4 +
z2∂∂z2 ∧ ∂∂z5 +
z3∂∂z3 ∧ ∂∂z6
+ ez4 − z5 ∂∂z1 ∧ ∂∂z2 +
ez5 − z6 ∂∂z2 ∧ ∂∂z3 +
∂∂z4 ∧ ∂∂z5 + ∂∂z5 ∧ ∂∂z6
(63)
Volume multivector fields
Ŵk ∧ Wn − k can be calculated in the manner
similar to R4 case and give rise to the well known conservation laws of
three particle Toda chain.
Y(1) = 16(z1 + z2 + z3) =
Ŵ ∧ W ∧ WW ∧ W ∧ W
Y(2) = 13
(z1z2 + z1z3 + z2z3
− ez4 − z5 − ez5 − z6)
= Ŵ ∧ Ŵ ∧ WW ∧ W ∧ W
Y(3) = z1z2z3 −
z3ez4 − z5 −
z1ez5 − z6
= Ŵ ∧ Ŵ ∧ ŴW ∧ W ∧ W
(64)
4. Non-Liouville symmetries
Besides Hamiltonian dynamical systems that admit invariant symplectic form
ω, there are dynamical systems that either are not Hamiltonian or
admit Hamiltonian realization but explicit form of symplectic structure ω
is unknown or too complex. However usually such a dynamical systems possess invariant volume form
Ω which like symplectic form can be effectively used in construction of
conservation laws. Note that volume form for given manifold is arbitrary differential form
of maximal degree (equal to the dimension of manifold).
In case of regular Hamiltonian systems, n-th outer power of the symplectic form ω
naturally gives rise to the invariant volume form known as Liouville form
Ω = ωn
(65)
and sometimes it is easier to work with Ω rather then with symplectic form itself.
In generic Liouville dynamical system time evolution is governed by equations of motion
ddtf = X(f)
(66)
where X is some smooth vector field that preserves Liouville volume form
Ω
ddtΩ = LXΩ = 0
(67)
Symmetry of equations of motion still can be defined by condition
ddt gz(f)
= gz(ddtf)
(68)
that in terms of vector fields implies that generator of symmetry E should
commute with time evolution operator X
[E , X] = 0
(69)
Throughout this chapter symmetry will be called non-Liouville if it is not conformal symmetry
of Ω, or in other words if
LEΩ ≠ cΩ
(70)
for any constant c.
Such a symmetries may be considered as analog of non-Noether symmetries
defined in Hamiltonian systems and similarly to the Hamiltonian case one can try
to construct conservation laws by means of generator of symmetry E
and invariant differential form Ω. Namely we have the following
theorem, which is reformulation of Hojman's theorem in terms of Liouville volume form.
Theorem 2.
Let (M, X, Ω) be Liouville dynamical system on the smooth
manifold M. Then, if the vector field E generates
non-Liouville symmetry, the function
J = LEΩΩ
(71)
is conservation law.
Proof.
By the definition
LEΩ = JΩ.
(72)
and J is not just constant
(again definition is correct since the space of volume forms is one dimensional).
By taking Lie derivative of this expression along the vector field X that
defines time evolution we get
LXLEΩ = L[X , E]Ω + LELXΩ
= LX(JΩ) = (LXJ)Ω + JLXΩ
(73)
but since Liouville volume form is invariant LXΩ = 0 and
vector field E is generator of symmetry satisfying [E , X] = 0
commutation relation we obtain
(LXJ)Ω = 0
(74)
or
ddtJ = LXJ = 0
(75)
Remark 3.
In fact theorem is valid for larger class of symmetries. Namely one can consider
symmetries with time dependent generators. Note however that in this case condition
[E , X] = 0 should be replaced by
∂∂tE = [E , X]
(76)
Note also that by calculating Lie derivative of conservation law J along
generator of the symmetry E one can recover additional conservation laws
J(m) = (LE)mΩ
(77)
Example 2.
Let us consider symmetry of three particle non periodic Toda chain. This dynamical system
with equations of motion
ż4 = z1
ż5 = z2
ż6 = z3
(78)
ż1 = − ez4 − z5
ż2 = ez4 − z5 − ez5 − z6
ż3 = ez5 − z6
(79)
possesses invariant volume form
Ω = dz1 ∧ dz2 ∧ dz3 ∧
dz4 ∧ dz5 ∧ dz6
(80)
The symmetry (61) is clearly non-Liouville one as far as
LEΩ = (z1 + z2 + z3) Ω
(81)
and main conservation law associated with this symmetry via Theorem 2 is total momentum
J = LEΩΩ = z1 + z2 + z3
(82)
Other conservation laws can be recovered by taking Lie derivative of J
along generator of symmetry E, in particular
J(1) = LEJ =
12z12 +
12z22 +
12z32 +
ez4 − z5 +
ez5 − z6
J(2) = LEJ(1) =
12 (z13 + z23 + z33)
+ 32 (z1 + z2)ez4 − z5 +
32 (z2 + z3)ez5 − z6
(83)
5. Lax Pairs
Presence of the non-Noether symmetry not only leads to a sequence of conservation laws, but also
endows the phase space with a number of interesting geometric structures and it appears that such a
symmetry is related to many important concepts used in theory of dynamical systems.
One of the such concepts is Lax pair that plays quite important role in construction
of completely integrable models.
Let us recall that Lax pair of Hamiltonian system on Poisson manifold M is
a pair (L , P) of smooth functions on M with values in some
Lie algebra g such that the time evolution of L is given by
adjoint action
ddtL = [L , P] = − adPL
(84)
where [ , ] is a Lie bracket on g. It is well known that each Lax
pair leads to a number of conservation laws. When g is some matrix Lie algebra
the conservation laws are just traces of powers of L
I(k) =
12
Tr(Lk)
(85)
since trace is invariant under coadjoint action
ddtI(k) = 12 ddt Tr(Lk) =
12 Tr(ddtLk)
= k2 Tr(Lk − 1ddtL)
= k2 Tr(Lk − 1[L , P]) = 12 Tr([Lk, P]) = 0
(86)
It is remarkable that each generator of the non-Noether
symmetry canonically leads to the Lax pair of a certain type.
Such a Lax pairs have definite geometric origin, their Lax matrices are formed
by coefficients of invariant tangent valued 1-form on the phase space.
In the local coordinates zs, where the bivector field
W, symplectic form ω and the generator
of the symmetry E have the following form
W = ∑rs Wrs∂∂zr ∧ ∂∂zr
ω = ∑rs ωrsdzr ∧ dzs
E = ∑s Es∂∂zs
(87)
corresponding Lax pair can be calculated explicitly.
Namely we have the following theorem (see also [55]-[56]):
Theorem 3.
Let (M , h) be regular Hamiltonian system on the 2n-dimensional
Poisson manifold M.
Then, if the vector field E on M generates the non-Noether symmetry,
the following 2n×2n matrix valued functions on M
Lab = ∑dc ωad
[
Ec∂Wdb∂zc −
Wbc∂Ed∂zc
+ Wdc∂Eb∂zc
]
Pab = ∑c
[
∂Wbc∂za·∂h∂zc + Wbc∂2h∂za∂zc
]
(88)
form the Lax pair (84) of the dynamical system (M , h).
Proof.
Let us consider the following operator on a space of 1-forms
ŔE(u) = Φω([E , ΦW(u)]) − LEu
(89)
(here ΦW and Φω
are maps induced by Poisson bivector field and symplectic form).
It is remarkable that ŔE appears to be invariant linear operator.
First of all let us show that ŔE is really linear,
or in other words, that for arbitrary 1-forms u and v
and function f operator ŔE has the following properties
ŔE(u + v) = ŔE(u) + ŔE(v)
(90)
and
ŔE(fu) = fŔE(u)
(91)
First property is obvious consequence of linearity of Schouten bracket, Lie derivative and
maps ΦW, Φω.
Second property can be checked directly
ŔE(fu) = Φω([E , ΦW(fu)]) − LE(fu)
= Φω([E , fΦW(u)]) − (LEf)u − fLEu
= Φω((LEf)ΦW(u))
+ Φω(f[E , ΦW(u)]) − (LEf)u − fLEu
= LEfΦωΦW(u) + fΦω([E , ΦW(u)])
− (LEf)u − fLEu
= f(Φω([E , ΦW(u)]) − LEu) = fŔE(u)
(92)
as far as ΦωΦW(u) = u.
Now let us check that ŔE is invariant operator
ddtŔE = LXhŔE =
LXh(ΦωLEΦW − LE)
= ΦωL[Xh , E]ΦW
− L[Xh, E] = 0
(93)
because, being Hamiltonian vector field, Xh commutes with maps
ΦW, Φω
(this is consequence of Liouville theorem) and commutes with E
as far as E generates the symmetry [Xh, E] = 0.
In the terms of the local coordinates ŔE has the following form
ŔE =
∑ab
Lab dza ⊗ ∂∂zb
(94)
and the invariance condition
ddtŔE = LW(h)ŔE = 0
(95)
yields
ddtŔE = ddt∑ab
Lab dza ⊗ ∂∂zb
= ∑ab (ddtLab) dza ⊗ ∂∂zb +
∑ab Lab (LW(h)dza) ⊗ ∂∂zb
+ ∑ab Lab dza ⊗ (LW(h)∂∂zb) =
∑ab (ddtLab) dza ⊗ ∂∂zb
+ ∑abcd
Lab∂Wad∂zc·∂h∂zd
dzc ⊗ ∂∂zb
+
∑abcd
LabWad∂2h∂zc∂zd
dzc ⊗ ∂∂zb
+
∑abcdLab∂Wcd∂zb·∂h∂zd
dza ⊗ ∂∂zc +
∑abcd LabWcd ∂2h∂zb∂zd
dza ⊗ ∂∂zc
= ∑ab
[
ddtLab +
∑c(PacLcb
− LacPcb)
]
dza ⊗ ∂∂zb = 0
(96)
or in matrix notations
ddtL = [L , P].
(97)
So, we have proved that the non-Noether symmetry canonically yields a Lax pair
on the algebra of linear operators on cotangent bundle over the phase space.
Remark 4. The conservation laws (85)
associated with the Lax pair (88) can be expressed in terms of the
integrals of motion ci in quite simple way:
I(k) = 12 Tr(Lk) = ∑s csk
(98)
This correspondence follows from the equation (40)
and the definition of the operator ŔE (89).
One can also write down recursion relation that determines conservation laws
I(k) in terms of conservation laws C(k)
I(m) + (− 1)mmC(m) +
m − 1∑k = 1
(− 1)k I(m − k)C(k) = 0
(99)
Example 3.
Let us calculate Lax matrix of two particle Toda chain
associated with non-Noether symmetry (53).
Using (88) it is easy to check that Lax matrix has eight nonzero elements
L =
z1
0
0
− ez3 − z4
0
z2
ez3 − z4
0
0
1
z1
0
− 1
0
0
z2
(100)
while matrix P involved in Lax pair
ddtL = [L , P]
(101)
has the following form
P =
0
0
1
0
0
0
0
1
− ez3 − z4
ez3 − z4
0
0
ez3 − z4
− ez3 − z4
0
0
(102)
The conservation laws associated with this Lax pair
are total momentum and energy of two particle Toda chain
I(1) = 12 Tr(L) = z1 + z2
I(2) = 12 Tr(L2) =
z12 + z22 + 2ez3 − z4
(103)
Similarly one can construct Lax matrix of three particle Toda chain, it has 16 nonzero elements
L =
z1
0
0
0
− ez4 − z5
0
0
z2
0
ez4 − z5
0
− ez5 − z6
0
0
z3
0
ez5 − z6
0
0
− 1
− 1
z1
0
0
1
0
− 1
0
z2
0
1
1
0
0
0
z3
(104)
with non-zero elements matrix P listed below
P =
0
0
0
1
0
0
0
0
0
0
1
0
0
0
0
0
0
1
− ez4 − z5
ez4 − z5
0
0
0
0
ez4 − z5
− ez4 − z5 − ez5 − z6
ez5 − z6
0
0
0
0
ez5 − z6
− ez5 − z6
0
0
0
(105)
Corresponding conservation laws reproduce total momentum, energy and second
Hamiltonian involved in bi-Hamiltonian realization of Toda chain
I(1) = 12 Tr(L) = z1 + z2
I(2) = 12 Tr(L2) =
z12 + z22 + z32 +
2ez4 − z5 + 2ez5 − z6
I(3) = 12 Tr(L3) =
z13 + z23 + z33 +
3(z1 + z2)ez4 − z5 +
3(z2 + z3)ez5 − z6
(106)
6. Involutivity of conservation laws
Now let us focus on the integrability issues. We know that
n integrals of motion are associated with each generator of non-Noether
symmetry, in the same time we know that, according to the Liouville-Arnold theorem,
regular Hamiltonian system (M, h) on 2n dimensional symplectic manifold
M is completely integrable (can be solved completely) if it admits
n functionally independent integrals of motion in involution.
One can understand functional independence of set of conservation laws
c1, c2 ... cn as
linear independence of either differentials of conservation laws
dc1, dc2 ... dcn or
corresponding Hamiltonian vector fields
Xc1, Xc2 ... Xcn.
Strictly speaking we can say that conservation laws c1, c2 ... cn
are functionally independent if Lesbegue measure of the set of points of phase space M
where differentials dc1, dc2 ... dcn become linearly dependent
is zero. Involutivity of conservation laws means that all possible Poisson brackets of
these conservation laws vanish pair wise
{ci , cj} = 0 i, j = 1... n
(107)
In terms of the vector fields, existence of involutive family of n
functionally independent conservation laws
c1, c2 ... cn
implies that corresponding Hamiltonian vector fields
Xc1, Xc2 ... Xcn
span Lagrangian subspace (isotropic subspace of dimension n)
of tangent space (at each point of M).
Indeed, due to property (23)
{ci , cj} = ω(Xci , Xcj) = 0
(108)
thus space spanned by Xc1, Xc2 ... Xcn
is isotropic. Dimension of this space is n so it is Lagrangian. Note also that distribution
Xc1, Xc2 ... Xcn
is integrable since due to (22)
[Xci , Xcj] = X{ci , cj} = 0
(109)
and according to Frobenius theorem there exists submanifold of M such that
distribution Xc1, Xc2 ... Xcn spans tangent
space of this submanifold. Thus for phase space geometry existence of complete involutive set
of integrals of motion implies existence of invariant Lagrangian submanifold.
Now let us look at conservation laws Y(1), Y(2) ... Y(n)
associated with generator of non-Noether symmetry. Generally speaking these conservation laws
might appear to be neither functionally independent nor involutive.
However it is reasonable to ask the question – what condition should be satisfied
by the generator of the non-Noether symmetry to ensure the involutivity
({Y(k) , Y(m)} = 0) of conserved quantities?
In Lax theory situation is very similar — each Lax matrix leads to the set of
conservation laws but in general this set is not involutive, however in Lax theory
there is certain condition known as Classical Yang-Baxter Equation (CYBE)
that being satisfied by Lax matrix ensures that conservation laws are in involution.
Since involutivity of the conservation laws is closely related to the integrability,
it is essential to have some analog of CYBE for the generator
of non-Noether symmetry. To address this issue we would like to propose the following theorem.
Theorem 4.
If the vector field E on 2n-dimensional
Poisson manifold M satisfies the condition
[[E[E , W]]W] = 0
(110)
and W bivector field has maximal rank (Wn ≠ 0)
then the functions (32) are in involution
{Y(k) , Y(m)} = 0
(111)
Proof. First of all let us note that
the identity (15) satisfied by the Poisson
bivector field W is responsible for the Liouville theorem
[W , W] = 0 ⇔ LW(f)W = [W(f) , W] = 0
(112)
that follows from the graded Jacoby identity satisfied by Schouten bracket.
By taking the Lie derivative of the expression (15)
we obtain another useful identity
LE[W , W] = [E[W , W]] = [[E , W] W] + [W[E , W]]
= 2[Ŵ , W] = 0.
(113)
This identity gives rise to the following relation
[Ŵ , W] = 0 ⇔ [Ŵ(f) , W] = − [Ŵ , W(f)]
(114)
and finally condition (110) ensures third identity
[Ŵ , Ŵ] = 0
(115)
yielding Liouville theorem for Ŵ
[Ŵ , Ŵ] = 0 ⇔ [Ŵ(f) , Ŵ] = 0
(116)
Indeed
[Ŵ , Ŵ] = [[E , W]Ŵ] = [[Ŵ , E]W]
= − [[E , Ŵ]W] = − [[E[E , W]]W] = 0
(117)
Now let us consider two different solutions ci ≠ cj
of the equation (40). By taking the Lie derivative of the equation
(Ŵ − ciW)n = 0
(118)
along the vector fields W(cj) and
Ŵ(cj) and using Liouville theorem for
W and Ŵ bivectors we obtain the following relations
(Ŵ −
ciW)n − 1(LW(cj)Ŵ
− {cj , ci}W) =
0,
(119)
and
(Ŵ −
ciW)n − 1(ciLŴ(cj)W
+ {cj , ci}∗W) = 0,
(120)
where
{ci , cj}∗ =
Ŵ(dci ∧ dcj)
(121)
is the Poisson bracket calculated by means of the bivector field Ŵ.
Now multiplying (119) by ci subtracting (120) and using
identity (114) gives rise to
({ci , cj}∗ −
ci{ci , cj})(Ŵ −
ciW)n − 1W = 0
(122)
Thus, either
{ci , cj}∗ −
ci{ci , cj} = 0
(123)
or the volume field
(Ŵ − ciW)n − 1W
vanishes. In the second case we can repeat
(119)-(122) procedure for
the volume field
(Ŵ − ciW)n − 1W
yielding after n
iterations Wn = 0 that according to our
assumption (that the dynamical system is regular) is not true.
As a result we arrived at (123) and by the simple
interchange of indices i ↔ j we get
{ci , cj}∗ −
cj{ci , cj} = 0
(124)
Finally by comparing (123) and (124) we obtain that
the functions ci are in involution with respect to the both
Poisson structures (since ci ≠ cj)
{ci , cj}∗ =
{ci , cj} = 0
(125)
and according to (41) the same is true for the integrals of motion
Y(k).
Remark 5. Theorem 4 is useful in multidimensional dynamical systems where involutivity of
conservation laws can not be checked directly.
7. Bi-Hamiltonian systems
Further we will focus on non-Noether symmetries that satisfy condition (110). Besides
yielding involutive families of conservation laws, such a symmetries appear to be related
to many known geometric structures such as bi-Hamiltonian systems [53]
and Frölicher-Nijenhuis operators (torsionless tangent valued differential 1-forms).
The relationship between non-Noether symmetries and bi-Hamiltonian structures was
already implicitly outlined in the proof of Theorem 4. Now let us pay more attention to
this issue.
Originally bi-Hamiltonian structures were introduced by F. Magri in analisys of
integrable infinite dimensional Hamiltonian systems such as Korteweg-de Vries (KdV) and
modified Korteweg-de Vries (mKdV) hierarchies, Nonlinear Schrödinger equation
and Harry Dym equation. Since that time bi-Hamiltonian formalism is effectively used
in construction of involutive families of conservation laws in integrable models
Generic bi-Hamiltonian structure on 2n dimensional manifold consists out
of two Poisson bivector fields W and Ŵ satisfying certain
compatibility condition [Ŵ , W] = 0. If, in addition, one of these bivector
fields is nondegenerate (Wn ≠ 0) then bi-Hamiltonian system
is called regular. Further we will discuss only regular bi-Hamiltonian systems.
Note that each Poisson bivector field by definition satisfies condition (15). So we actually
impose four restrictions on bivector fields W and Ŵ
[W , W] = [Ŵ , W] = [Ŵ , Ŵ] = 0
(126)
and
Wn ≠ 0
(127)
During the proof of Theorem 4 we already showed that bivector fields
W and Ŵ = [E , W] satisfy conditions (126)
(see (112)-(116)),
thus we can formulate the following statement
Theorem 5.
Let (M , h) be regular Hamiltonian system on the 2n-dimensional
manifold M endowed with regular Poisson bivector field W.
Then, if the vector field E on M generates the non-Noether symmetry,
and satisfies condition
[[E[E , W]]W] = 0,
(128)
the following bivector fields on M
W, Ŵ = [E , W]
(129)
form invariant bi-Hamiltonian system
([W , W] = [Ŵ , W] = [Ŵ , Ŵ] = 0).
Proof. See proof of Theorem 4.
Remark 6. Bi-Hamiltonian systems obtained by taking Lie derivative of Poisson bivector
along some vector field were studied in [70]
Example 4. One can check that the non-Noether symmetry (53) satisfies
condition (110) while bivector fields
W =
∂∂z1 ∧ ∂∂z3 +
∂∂z2 ∧ ∂∂z4
(130)
and
Ŵ = [E , W] = z1∂∂z1 ∧ ∂∂z3
+ z2∂∂z2 ∧ ∂∂z4
+ ez3 − z4 ∂∂z1 ∧ ∂∂z2 +
∂∂z3 ∧ ∂∂z4
(131)
form bi-Hamiltonian system [W , W] = [W , Ŵ] = [Ŵ , Ŵ] = 0.
Similarly, one can recover bi-Hamiltonian system of three particle Toda chain associated
with symmetry (61). It is formed by bivector fields
W = ∂∂z1 ∧ ∂∂z4 + ∂∂z2 ∧ ∂∂z5 +
∂∂z3 ∧ ∂∂z6
(132)
and
Ŵ = [E , W] = z1∂∂z1 ∧ ∂∂z4 +
z2∂∂z2 ∧ ∂∂z5 +
z3∂∂z3 ∧ ∂∂z6
+ ez4 − z5 ∂∂z1 ∧ ∂∂z2 +
ez5 − z6 ∂∂z2 ∧ ∂∂z3 +
∂∂z4 ∧ ∂∂z5 + ∂∂z5 ∧ ∂∂z6
(133)
In terms of differential forms bi-Hamiltonian structure is formed by couple of
closed differential 2-forms: symplectic form ω
(such that dω = 0 and ωn ≠ 0)
and ω∗ = LEω
(clearly dω∗ = dLEω
= LEdω = 0). It is important that by taking Lie derivative of
Hamilton's equation
iXhω + dh = 0
(134)
along the generator E of symmetry
LE(iXhω + dh) =
i[E , Xh]ω + iXhLEω + LEdh =
iXhω∗ + dLEh =
0
(135)
one obtains another Hamilton's equation
iXhω∗ + dh∗ = 0
(136)
where h∗ = LEh. This is actually second Hamiltonian realization
of equations of motion and thus under certain conditions existence of non-Noether symmetry
gives rise to additional presymplectic structure ω∗
and additional Hamiltonian realization of the dynamical system.
In many integrable models admitting bi-Hamiltonian realization (including Toda chain,
Korteweg-de Vries hierarchy, Nonlinear Schrödinger equation, Broer-Kaup system and
Benney system) non-Noether symmetries that are responsible for existence of bi-Hamiltonian structures
has been found and motivated further investigation of relationship between
symmetries and bi-Hamiltonian structures. Namely it seems to be interesting to know
whether in general case existence of bi-Hamiltonian structure is related to non-Noether symmetry.
Let us consider more general case and suppose that we have couple of differential 2-forms
ω and ω∗
such that
dω = dω∗ = 0, ωn ≠ 0
(137)
iXhω + dh = 0
(138)
and
iXhω∗ + dh∗ = 0
(139)
The question is whether there exists vector field E (generator of non-Noether symmetry)
such that [E , Xh] = 0 and
ω∗ = LEω.
The answer depends on ω∗.
Namely if ω∗ is exact form
(there exists 1-form θ∗ such that
ω∗ = dθ∗)
then one can argue that such a vector field exists and thus any
exact bi-Hamiltonian structure is related to hidden non-Noether
symmetry. To outline proof of this statement let us introduce
vector field E∗ defined by
iE∗ω = θ∗
(140)
(such a vector field always exist because ω
is nondegenerate 2-form).
By construction
LE∗ ω = ω∗
(141)
Indeed
LE∗ω = diE∗ω +
iE∗dω = dθ∗ = ω∗
(142)
And
i[E∗, Xh]ω =
LE∗(iXhω)
− iXhLE∗ω =
− d(E∗(h)
− h∗) = − dh'
(143)
In other words [Xh , E∗] is Hamiltonian vector field
[Xh , E] = Xh'
(144)
One can also construct locally Hamiltonian vector field Xg,
that satisfies the same commutation relation. Namely let us define
function (in general case this can be done only locally)
g(z) = t∫0 h'dt
(145)
where integration along solution of Hamilton's equation, with fixed origin and end point in
z(t) = z, is assumed.
And then it is easy to verify that locally Hamiltonian vector field associated with g(z),
by construction, satisfies the same commutation relations as
E∗ (namely [Xh , Xg] = Xh').
Using E∗ and Xh'
one can construct generator of non-Noether symmetry —
non-Hamiltonian vector field E = E∗ − Xg
commuting with Xh and satisfying
LEω = LE∗ω
− LXgω = LE∗ω = ω∗
(146)
(thanks to Liouville's theorem LXgω = 0). So in
case of regular Hamiltonian system every exact bi-Hamiltonian structure is
naturally associated with some (non-Noether) symmetry of space of solutions.
In case where bi-Hamiltonian structure is not exact
(ω∗ is closed but not exact) then due to
ω∗ = LEω =
diEω + iEdω = diEω
(147)
it is clear that such a bi-Hamiltonian system is not related to symmetry.
However in all known cases bi-Hamiltonian structures seem to be exact.
8. Bidifferential calculi
Another important concept that is often used in theory of dynamical systems and may
be related to the non-Noether symmetry is the bidifferential calculus (bicomplex approach).
Recently A. Dimakis and F. Müller-Hoissen
applied bidifferential calculi to the wide range of integrable models
including KdV hierarchy, KP equation, self-dual Yang-Mills equation,
Sine-Gordon equation, Toda models, non-linear Schrödinger
and Liouville equations. It turns out that these models can be effectively
described and analyzed using the bidifferential calculi
[17], [24].
Here we would like to show that each generator of non-Noether symmetry
satisfying condition [[E[E , W]]W] = 0 gives rise to certain
bidifferential calculus.
Before we proceed let us specify what kind of bidifferential calculi we plan to consider.
Under the bidifferential calculus we mean the graded algebra of differential forms
over the phase space
Ω =
∞∪k = 0
Ω(k)
(148)
(Ω(k) denotes the space of k-degree differential forms)
equipped with a couple of differential operators
d, đ : Ω(k) → Ω(k + 1)
(149)
satisfying conditions
d2 = đ2 = dđ + đd = 0
(see [24]). In other words we have two De Rham
complexes M, Ω, d and M, Ω, đ
on algebra of differential forms over the phase space. And these complexes satisfy
certain compatibility condition — their differentials anticommute with each other
dđ + đd = 0.
Now let us focus on non-Noether symmetries.
It is interesting that if generator of the non-Noether symmetry satisfies
equation [[E[E , W]]W] = 0 then we are able to construct an invariant
bidifferential calculus of a certain type.
This construction is summarized in the following theorem:
Theorem 6.
Let (M , h) be regular Hamiltonian system on the Poisson manifold M.
Then, if the vector field E on M generates the non-Noether symmetry
and satisfies the equation
[[E[E , W]]W] = 0,
(150)
the differential operators
du =
Φω([W , ΦW(u)])
(151)
đu =
Φω([[E , W]ΦW(u)])
(152)
form invariant bidifferential calculus
(d2 = đ2 = dđ + đd = 0)
over the graded algebra of differential forms on M.
Proof. First of all we have to show that d and đ
are really differential operators , i.e., they are linear maps from
Ω(k) into
Ω(k + 1), satisfy derivation property and
are nilpotent (d2 = đ2 = 0).
Linearity is obvious and follows from the linearity of the Schouten bracket [ , ]
and ΦW, Φω
maps. Then, if u is a k-degree form
ΦW maps it on k-degree multivector field and
the Schouten brackets [W , ΦW(u)] and
[[E , W]ΦW(u)] result the
k + 1-degree multivector fields that are mapped on k + 1-degree
differential forms by Φω.
So, d and đ
are linear maps from Ω(k) into
Ω(k + 1).
Derivation property follows from the same feature of the Schouten bracket
[ , ] and linearity of
ΦW and
Φω maps.
Now we have to prove the nilpotency of d and đ.
Let us consider d2u
d2u =
Φω([W , ΦW(Φω([W , ΦW(u)]))])
= Φω([W[W , ΦW(u)]]) = 0
(153)
as a result of the property (112) and the Jacoby identity for [ , ] bracket.
In the same manner
đ2u =
Φω([[W , E][[W , E]ΦW(u)]]) = 0
(154)
according to the property (116) of
[W , E] = Ŵ and the Jacoby identity.
Thus, we have proved that d and đ are differential operators
(in fact d is ordinary exterior differential and the expression
(151) is its well known representation in terms of Poisson bivector field).
It remains to show that the compatibility condition dđ + đd = 0
is fulfilled. Using definitions of d, đ and the Jacoby identity we get
(dđ + đd)(u) =
Φω([[[W , E]W]ΦW(u)]) = 0
(155)
as far as (114) is satisfied.
So, d and đ form the bidifferential calculus over the graded
algebra of differential forms.
It is also clear that the bidifferential calculus d, đ
is invariant, since both d and đ commute with time evolution
operator W(h) = {h, }.
Remark 7.
Conservation laws that are associated with the bidifferential calculus
(151) (152)
and form Lenard scheme (see [24]):
(k + 1)đI(k) = kdI(k + 1)
(156)
coincide with the sequence of integrals of motion (98).
Proof of this correspondence lays outside the scope of present manuscript,
but can be done in the manner similar to [17].
Example 5.
The symmetry (53) endows R4 with bicomplex structure
d, đ where d is ordinary exterior derivative while đ
is defined by
đz1 = z1dz1 − ez3 − z4dz4
đz2 = z2dz2 + ez3 − z4dz3
đz3 = z1dz3 + dz2
đz4 = z2dz4 − dz1
(157)
and is extended to whole De Rham complex by linearity, derivation property and
compatibility property dđ + đd = 0.
By direct calculations one can verify that calculus constructed in this way
is consistent and satisfies đ2 = 0 property.
To illustrate technique let us explicitly check that đ2z1 = 0.
Indeed
đ2z1 = đđz1 =
đ(z1dz1 − ez3 − z4dz4)
= đz1 ∧ dz1 + z1đdz1
− ez3 − z4đz3 ∧ dz4
+ ez3 − z4đz4 ∧ dz4
− ez3 − z4đdz4
= đz1 ∧ dz1 − z1dđz1
− ez3 − z4đz3 ∧ dz4
+ ez3 − z4đz4 ∧ dz4
+ ez3 − z4dđz4 = 0
(158)
Because of properties
đz1 ∧ dz1 =
ez3 − z4dz1 ∧ dz4,
(159)
− z1dđz1 =
z1ez3 − z4dz3 ∧ dz4,
(160)
− ez3 − z4đz3 ∧ dz4
= − z1ez3 − z4dz1 ∧ dz4
− ez3 − z4dz2 ∧ dz4,
(161)
ez3 − z4đz4 ∧ dz4 =
ez3 − z4dz2 ∧ dz4
(162)
and
ez3 − z4dđz4 =
− ez3 − z4dz1 ∧ dz4
(163)
Similarly one can show that
đ2z2 = đ2z3 = đ2z4 = 0
(164)
and thus đ is nilpotent operator đ2 = 0.
Note also that conservation laws
I(1) = z1 + z2
I(2) = z12 + z22
+ 2ez3 − z4
(165)
form the simplest Lenard scheme
2đI(1) = dI(2)
(166)
Similarly one can construct bidifferential calculus associated with non-Noether
symmetry (61) of three particle Toda chain. In this case đ
can be defined by
đz1 = z1dz1 − ez4 − z5dz5
đz2 = z2dz2 + ez4 − z5dz4
− ez5 − z6dz6
đz3 = z3dz3 + ez5 − z6dz5
đz4 = z1dz4 − dz2 − dz3
đz5 = z2dz5 + dz1 − dz3
đz6 = z3dz6 + dz1 + dz2
(167)
and as in case of two particle Toda it
can be extended to whole De Rham complex by linearity, derivation property and
compatibility property dđ + đd = 0.
One can check that conservation laws of Toda chain
I(1) = z1 + z2
I(2) =
z12 + z22 + z32 +
2ez4 − z5 + 2ez5 − z6
I(3) =
z13 + z23 + z33 +
3(z1 + z2)ez4 − z5 +
3(z2 + z3)ez5 − z6
(168)
form Lenard scheme
2đI(1) = dI(2)
(169)
3đI(2) = 2dI(3)
(170)
9. Frölicher-Nijenhuis geometry
Finally we would like to reveal some features of the operator
ŔE
(89) and to show how Frölicher-Nijenhuis geometry arises in
Hamiltonian system that possesses certain non-Noether symmetry.
From the geometric properties of the tangent valued forms we know
that the traces of powers of a linear operator F
on tangent bundle are in involution whenever its Frölicher-Nijenhuis torsion
T(F) vanishes, i. e. whenever for arbitrary vector fields X,Y the condition
T(F)(X , Y) = [FX , FY] −
F([FX , Y] + [X , FY] − F[X , Y]) = 0
(171)
is satisfied.
Torsionless forms are also called Frölicher-Nijenhuis operators and are widely used in
theory of integrable models, where they play role of recursion operators and are used
in construction of involutive family of conservation laws.
We would like to show that each generator of non-Noether symmetry satisfying equation
[[E[E , W]]W] = 0
canonically leads to invariant Frölicher-Nijenhuis operator on tangent
bundle over the phase space. This operator can be expressed in terms of generator of symmetry
and isomorphism defined by Poisson bivector field. Strictly speaking we have the following theorem.
Theorem 7.
Let (M , h) be regular Hamiltonian system on the Poisson manifold M.
If the vector field E on M generates the non-Noether symmetry
and satisfies the equation
[[E[E , W]]W] = 0
(172)
then the linear operator, defined for
every vector field X by equation
RE(X) =
ΦW(LEΦω(X))
− [E , X]
(173)
is invariant Frölicher-Nijenhuis operator on M.
Proof.
Invariance of RE follows from the invariance of the
ŔE defined by (89)
(note that for arbitrary 1-form vector field u and vector field X
contraction iXu has the property
iREXu =
iXŔEu,
so RE is actually transposed to
ŔE).
It remains to show that the condition (110) ensures vanishing of the
Frölicher-Nijenhuis torsion T(RE) of
RE, i.e. for arbitrary vector fields X, Y we must get
T(RE)(X , Y) = [RE(X) , RE(Y)] −
RE([RE(X) , Y]
+ [X , RE(Y)] − RE([X , Y])) = 0
(174)
First let us introduce the following auxiliary 2-forms
ω = Φω(W), ω∗ = ŔEω ω∗∗ = ŔEω∗
(175)
Using the realization (151) of the differential d
and the property (15) yields
dω = Φω([W , W]) = 0
(176)
Similarly, using the property (114) we obtain
dω∗ =
dΦω([E , W]) − dLEω =
Φω([[E , W]W]) −
LEdω = 0
(177)
And finally, taking into account that
ω∗ = 2Φω([E , W])
and using the condition (110), we get
dω∗∗ =
2Φω([[E[E , W]]W])
− 2dLEω∗ =
− 2LEdω∗ = 0
(178)
So the differential forms
ω, ω∗, ω∗∗
are closed
dω = dω∗ = dω∗∗ = 0
(179)
Now let us consider the contraction of T(RE) and ω.
iT(RE)(X , Y)ω =
i[REX , REY]ω −
i[REX , Y]ω∗ −
i[X , REY]ω∗ +
i[X , Y]ω∗∗
=LREXiYω∗ −
iREYLXω∗ −
LREXiYω∗ +
iYLREXω∗ −
LXiREYω∗ +
iREYLXω∗ +
i[X , Y]ω∗∗
= iYLXω∗∗ −
LXiYω∗∗ +
i[X , Y]ω∗∗ = 0
(180)
where we used (175) (179),
the property
LXiYω =
iYLXω + i[X , Y]ω
(181)
of the Lie derivative and the relations of the following type
LREXω =
diREXω + iREXdω
= diXω∗
= LXω∗ −
iXdω∗ = LXω∗
(182)
So we proved that for arbitrary vector fields X, Y
the contraction of T(RE)(X , Y) and ω vanishes.
But since W bivector is non-degenerate
(Wn ≠ 0), its counter image
ω = Φω(W)
(183)
is also non-degenerate and vanishing of the contraction (180)
implies that the torsion T(RE) itself is zero.
So we get
T(RE)(X , Y) = [RE(X) , RE(Y)] −
RE([RE(X) , Y]
+ [X , RE(Y)] − RE([X , Y])) = 0
(184)
Example 6. The operator RE associated with non-Noether
symmetry (53) reproduces well known Frölicher-Nijenhuis operator
RE =
z1dz1 ⊗ ∂∂z1 −
dz1 ⊗ ∂∂z4 +
z2dz2 ⊗ ∂∂z2 +
dz2 ⊗ ∂∂z3
+ z1dz3 ⊗ ∂∂z3 +
ez3 − z4dz3 ⊗ ∂∂z2 +
z2dz4 ⊗ ∂∂z4 −
ez3 − z4dz4 ⊗ ∂∂z1
(185)
(compare with [30]).
The operator ŔE
plays the role of recursion operator for conservation laws
I(1) = z1 + z2
I(2) = z12 + z22
+ 2ez3 − z4
(186)
Indeed one can check that
2ŔE(dI(1)) = dI(2)
(187)
Similarly using non-Noether symmetry (61) one can construct recursion operator of
three particle Toda chain
RE = z1dz1 ⊗ ∂∂z1
− ez4 − z5dz5 ⊗ ∂∂z1
+ z2dz2 ⊗ ∂∂z2 +
ez4 − z5dz4 ⊗ ∂∂z2
− ez5 − z6dz6 ⊗ ∂∂z2+
z3dz3 ⊗ ∂∂z3 +
ez5 − z6dz5 ⊗ ∂∂z3
+ z1dz4 ⊗ ∂∂z4
− dz2 ⊗ ∂∂z4
− dz3 ⊗ ∂∂z4
+ z2dz5 ⊗ ∂∂z5 + dz1 ⊗ ∂∂z5
− dz3 ⊗ ∂∂z5
+ z3dz6 ⊗ ∂∂z6 +
dz1 ⊗ ∂∂z6 + dz2 ⊗ ∂∂z6
(188)
and as in case of two particle Toda chain, operator ŔE
appears to be recursion operator for conservation laws
I(1) = z1 + z2
I(2) =
z12 + z22 + z32 +
2ez4 − z5 + 2ez5 − z6
I(3) =
z13 + z23 + z33
+ 3(z1 + z2)ez4 − z5
+ 3(z2 + z3)ez5 − z6
(189)
and fulfills the following recursion condition
dI(3) = 3ŔE(dI(2)) =
6(ŔE)2(dI(1))
(190)
10. One-parameter families of conservation laws
One-parameter group of transformations gz
defined by (28) naturally acts on algebra of integrals of motion.
Namely for each conservation law
ddtJ = 0
(191)
one can define one-parameter family of conserved quantities J(z)
by applying group of transformations gz to J
J(z) = gz(J) = ezLEJ =
J + zLEJ + ½(zLE)2J + ...
(192)
Property (29) ensures that J(z) is conserved for arbitrary values
of parameter z
ddtJ(z) =
ddtgz(J) =
gz
(ddtJ) = 0
(193)
and thus each conservation law gives rise to whole family of conserved
quantities that form orbit of group of transformations ga.
Such an orbit J(z) is called involutive if conservation laws that form
it are in involution
{J(z1) , J(z2)} = 0
(194)
(for arbitrary values of parameters z1, z2). On 2n dimensional
symplectic manifold each involutive family that contains n functionally independent
integrals of motion naturally gives rise to integrable system (due to Liouville-Arnold theorem).
So in order to identify those orbits that may be related to integrable models it
is important to know how involutivity of family of conserved quantities J(z)
is related to properties of initial conserved quantity J(0) = J and nature of
generator E of group gz = ezLE.
In other words we would like to know what condition must be satisfied by generator of
symmetry E and integral of motion J to ensure that
{J(z1) , J(z2)} = 0. To address this issue and to describe class of vector fields
that possess nontrivial involutive orbits we would like to propose the following
theorem
Theorem 8.
Let M be Poisson manifold endowed with 1-form s
such that
[W[W(s),W](s)] = c0[W(s)[W(s) ,W]] (c0 ≠ − 1)
(195)
Then each function J satisfying property
W(LW(s)dJ) = c1[W(s),W](dJ) (c1 ≠ 0)
(196)
(c0,1 are some constants) gives rise to involutive
set of functions
J(m) = (LW(s))mJ {J(m), J(k)} = 0
(197)
Proof.
First let us inroduce linear operator R on bundle of multivector fields and define it
for arbitrary multivector field V by condition
R(V) = ½ ([W(s),V] − ΦW(LW(s)Φω(V)))
(198)
Proof of linearity of this operator is identical to proof given for
(89) so we will skip it. Further it is clear that
R(W) = [W(s),W]
(199)
and
R2(W) = R([W(s),W]) = ½([W(s)[W(s),W]] − ΦW((LW(s))2ω))
= ½(1 + c0)[W(s)[W(s),W]]
(200)
where we used property
ΦW((LW(s))2ω) =
ΦW(LW(s)LW(s)ω)
= ΦW(iW(s)dLW(s)ω) +
ΦW(diW(s)LW(s)ω)
= [W,ΦW(iW(s)LW(s)ω)] =
[W[W(s),W](s)] = c0[W(s)[W(s),W]]
(201)
At the same time by taking Lie derivative of
(199) along the vector field W(s)
one gets
[W[W(s),W](s)] = (LW(s)R + R2)(W)
(202)
comparing (200) and (202) yields
(1 + c0)(LW(s)R + R2) = 2R2
(203)
and thus
(1 + c0)LW(s)R = (1 − c0)R2
(204)
Further let us rewrite condition (196) as follows
W(LW(s)dJ) = c1R(W)(dJ)
(205)
due to linearity of operator R this condition can be extended to
Rm(W)(LW(s)dJ) = c1Rm + 1(W)(dJ)
(206)
Now assuming that the following condition is true
W((LW(s))mdJ) = cmRm(W)(dJ)
(207)
let us take its Lie derivative along vector field W(s).
We get
R(W)((LW(s))mdJ) + W((LW(s))m + 1dJ)
= mcm1 − c01 + c0Rm + 1(W)(dJ) + cmRm + 1(W)(dJ)
(208)
where we used properties (199) and (204).
Note also that (207) together with linearity of operator R
imply that
RkW((LW(s))mdJ) = cmRk + m(W)(dJ)
(209)
and thus (208) reduces to
W((LW(s))m + 1dJ)
= cm + 1Rm + 1(W)(dJ)
(210)
where cm + 1 is defined by
(1 + c0)cm + 1
= mcn(1 − c0)
(211)
So we proved that if assumtion (207) is valid for m
then it is also valid for m + 1, we also know that for m = 1 it
matches (205) and thus by induction we proved that condition
(207) is valid for arbitrary m while cn
can be determined by
cm(1 + c0)m − 1 = c0(m − 1)!(1 − c0)m − 1
(212)
Now using (207) and (209)
it is easy to show that functions (LW(s))mJ are in involution.
Indeed
{(LW(s))mJ, (LW(s))kJ} =
W(d(LW(s))mJ ∧ d(LW(s))kJ)
= W((LW(s))mdJ ∧ (LW(s))kdJ) =
cmckW(dJ ∧ dJ) = 0
(213)
So we have proved that the functions (197) are in involution.
Further we will use this theorem to prove involutivity of family
of conservation laws constructed using non-Noether symmetry of Toda chain.
11. Toda Model
To illustrate features of non-Noether symmetries we often
refer to two and three particle non-periodic Toda systems.
However it turns out that non-Noether symmetries are present in
generic n-particle non-periodic Toda chains as well, moreover they preserve
basic features of symmetries (53), (61).
In case of n-particle Toda model symmetry yields n
functionally independent conservation laws in involution,
gives rise to bi-Hamiltonian structure of Toda hierarchy,
reproduces Lax pair of Toda system, endows phase space with
Frölicher-Nijenhuis operator and leads to invariant
bidifferential calculus on algebra of differential forms over phase space
of Toda system.
First of all let us remind that Toda model is
2n dimensional Hamiltonian system that describes the motion
of n particles on the line governed by the exponential interaction.
Equations of motion of the non periodic n-particle Toda model are
ddtqs = ps
ddtps = ε(s − 1)eqs − 1 − qs −
ε(n − s)eqs − qs + 1
(214)
(ε(k) = − ε(− k) = 1 for any natural
k and ε(0) = 0) and can be rewritten in Hamiltonian form
(24) with canonical Poisson bracket defined by Poisson bivector
W = n∑s = 1 ∂∂ps ∧ ∂∂qs
(215)
and Hamiltonian equal to
h = ½n∑s = 1ps2 +
n − 1∑s = 1eqs − qs + 1
(216)
Note that in two and three particle case we have used slightly different notations
zs = ps
zn + s = qs s = 1, 2, (3); n = 2(3)
(217)
for local coordinates.
The group of transformations gz generated by the vector field
E will be symmetry of Toda chain if for each
ps, qs satisfying Toda equations
(214)
gz(ps), gz(qs)
also satisfy it.
Substituting infinitesimal transformations
gz(ps) = ps + zE(ps) + O(z2)
gz(ps) = qs + zE(qs) + O(z2)
(218)
into (214) and grouping first order terms gives rise to the
conditions
ddtE(qs) = E(ps)
ddtE(ps) = ε(s − 1)eqs − 1 − qs
(E(qs − 1) − E(qs)) − ε(n − s)eqs − qs + 1
(E(qs) − E(qs + 1))
(219)
One can verify that the vector field defined by
E(ps) = ½ps2 +
ε(s − 1)(n − s + 2)eqs − 1 − qs −
ε(n − s)(n − s) eqs − qs + 1
+ t2(ε(s − 1)(ps − 1 + ps)
eqs − 1 − qs −
ε(n − s)(ps + ps + 1)
eqs − qs + 1)
E(qs) = (n − s + 1)ps −
½s − 1∑k = 1 pk
+ ½n∑k = s + 1 pk
+ t2(ps2 +
ε(s − 1)eqs − 1 − qs +
ε(n − s)eqs − qs + 1)
(220)
satisfies (31) and generates symmetry of Toda chain.
It appears that this symmetry is non-Noether since it does not
preserve Poisson bracket structure [E , W] ≠ 0
and additionally one can check that Yang-Baxter equation
[[E[E , W]]W] = 0 is satisfied.
This symmetry may play important role in
analysis of Toda model. First let us note that calculating LEW
leads to the following Poisson bivector field
Ŵ = [E , W] =
n∑s = 1 ps∂∂ps ∧ ∂∂qs
+ n − 1∑s = 1 eqs − qs + 1 ∂∂ps ∧ ∂∂qs + 1
+ ∑r > s ∂∂qs ∧ ∂∂qr
(221)
and together W and LEW give rise to
bi-Hamiltonian structure of Toda model (compare with [30]).
Thus bi-Hamiltonian realization of Toda chain can be considered as manifestation
of hidden symmetry.
In terms of bivector fields these bi-Hamiltonian system is formed by
The conservation laws (45) associated with the symmetry reproduce well known
set of conservation laws of Toda chain.
I(1) = C(1) = n∑s = 1ps
I(2) = (C(1))2 − 2C(2) =
n∑s = 1 ps2 +
2n − 1∑s = 1
eqs − qs + 1
I(3) = C(1))3 − 3C(1)C(2)
+ 3C(3) = n∑s = 1 ps3 +
3n − 1∑s = 1 (ps + ps + 1) eqs − qs + 1
I(4) = C(1))4 − 4(C(1))2C(2) +
2(C(2))2 + 4C(1)C(3) − 4C(4)
= n∑s = 1 ps4 +
4n − 1∑s = 1
(ps2 + 2psps + 1 + ps + 12)
eqs − qs + 1
+ 2n − 1∑s = 1 e2(qs − qs + 1) +
4n − 2∑s = 1 eqs − qs + 2
I(m) = (− 1)m + 1mC(m) +
m − 1∑k = 1
(− 1)k + 1I(m − k)C(k)
(222)
The condition [[E[E , W]]W] = 0 satisfied by generator of the
symmetry E ensures that the conservation laws are in involution
i. e. {C(k) , C(m)} = 0.
Thus the conservation laws as well as the bi-Hamiltonian structure
of the non periodic Toda chain appear to be associated with non-Noether symmetry.
Using formula (88) one can calculate Lax pair
associated with symmetry (220).
Lax matrix calculated in this way has the following non-zero entries
(note that in case of n = 2 and n = 3 this formula yields matrices
(102)-(105))
Lk, k = Ln + k, n + k = pk
Ln + k, k + 1 = − Ln + k + 1, k =
ε(n − k)eqk − qk + 1
Lk, n + m = ε(m − k)
m, k = 1, 2, ... , n
(223)
while non-zero entries of P matrix involved in Lax pair are
Pk, n + k = 1
Pn + k, k = − ε(k − 1)eqk − 1 − qk
− ε(n − k)eqk − qk + 1
Pn + k, k + 1 = ε(n − k)eqk − qk + 1
Pn + k, k − 1 = ε(k − 1)eqk − 1 − qk
k = 1, 2, ... , n
(224)
This Lax pair constructed from generator of non-Noether symmetry
exactly reproduces known Lax pair of Toda chain.
Like two and three particle Toda chain, n-particle Toda model also admits
invariant bidifferential calculus on algebra of differential forms over the phase space.
This bidifferential calculus can be constructed using non-Noether symmetry (see (152)),
it consists out of two differential operators d, đ
where d is ordinary exterior derivative while đ
can be defined by
đqs = psdqs + ∑r > sdpr
− ∑s > rdpr
đps = psdps − eqs − qs + 1dqs + 1
+ eqs − 1 − qsdqs
(225)
and is extended to whole De Rham complex by linearity, derivation property and
compatibility property dđ + đd = 0.
By direct calculations one can verify that calculus constructed in this way
is consistent and satisfies đ2 = 0 property.
One can also check that conservation laws (222) form Lenard scheme
(k + 1)đI(k) = kdI(k + 1)
(226)
Further let us focus on Frölicher-Nijenhuis geometry. Using formula (173)
one can construct invariant Frölicher-Nijenhuis operator, out of generator of non-Noether
symmetry of Toda chain. Operator constructed in this way has the form
ŔE = n∑s = 1ps(dps ⊗ ∂∂qs + dqs ⊗ ∂∂ps)
− n − 1∑s = 1 eqs − qs + 1dqs + 1 ⊗ ∂∂ps
+ n − 1∑s = 1 eqs − 1 − qsdqs ⊗ ∂∂ps
− ∑s > r (dps ⊗ ∂∂qr − dpr ⊗ ∂∂qs)
(227)
One can check that Frölicher-Nijenhuis torsion of this operator vanishes and
it plays role of recursion operator for n-particle Toda chain in sense that conservation laws
I(k) satisfy recursion relation
(k + 1)RE(dI(k)) = kdI(k + 1)
(228)
Thus non-Noether symmetry of Toda chain not only leads to
n functionally independent conservation laws in involution, but also
essentially enriches phase space geometry by endowing it with
invariant Frölicher-Nijenhuis operator, bi-Hamiltonian system,
bicomplex structure and Lax pair.
Finally, in order to outline possible applications of Theorem 8 let us study
action of non-Noether symmetry (220) on conserved quantities
of Toda chain. Vector field E defined by (220) generates
one-parameter group of transformations (28) that maps arbitrary
conserved quantity J to
J(z) = J + zJ(1) + z22!J(2) +
z33!J(3) + ⋯
(229)
where
J(m) = (LE)mJ
(230)
In particular let us focus on family of conserved quantities obtained by action of
ga = eaLE on total momenta of Toda chain
J = n∑s = 1 ps
(231)
By direct calculations one can check that family J(z), that forms orbit
of non-Noether symmetry generated by E, reproduces entire involutive
family of integrals of motion (222). Namely
J(1) = LEJ = ½ n∑s = 1
ps2 + n − 1∑s = 1
eqs − qs + 1
J(2) = LEJ(1) = (LE)2J =
12 n∑s = 1ps3 +
32n − 1∑s = 1 (ps + ps + 1)
eqs − qs + 1
J(3) = LEJ(2) = (LE)3J =
¾ n∑s = 1ps4 +
3n − 1∑s = 1(ps2 + 2psps + 1 +
ps + 12)eqs − qs + 1
+ 32 n − 1∑s = 1 e2(qs − qs + 1) +
3n − 2∑s = 1 eqs − qs + 2
J(m) = LEJ(m − 1) = (LE)mJ
(232)
Involutivity of this set of conservation laws can be verified using Theorem 8.
In particular one can notice that differential 1-form s defined by
E = W(s)
(233)
(where E is generator of non-Noether symmetry (220))
satisfies condition
[W[W(s),W](s)] = 3[W(s)[W(s) ,W]]
(234)
while conservation law J defined by (231)
has property
W(LW(s)dJ) = − [W(s),W](dJ)
(235)
and thus according to Theorem 8 conservation laws (232)
are in involution.
12. Korteweg-de Vries equation
Toda model provided good example of finite dimensional integrable Hamiltonian system
that possesses non-Noether symmetry. However there are many
infinite dimensional integrable Hamiltonian systems and in this case in
order to ensure integrability one should construct
infinite number of conservation laws. Fortunately in several integrable models
this task can be effectively simplified by identifying appropriate non-Noether symmetry.
First let us consider well known infinite dimensional integrable Hamiltonian system –
Korteweg-de Vries equation (KdV). The KdV equation has the following form
ut + uxxx + uux = 0
(236)
(here u is smooth function of (t, x) ∈ R2).
The generators of symmetries of KdV should satisfy condition
E(u)t + E(u)xxx +
uxE(u) + uE(u)x = 0
(237)
which is obtained by substituting infinitesimal transformation
u → u + zE(u) + O(z2) into KdV equation and grouping first order
terms.
Later we will focus on the symmetry generated by the following vector field
E(u) = 2uxx + 23u2 + 16uxv +
x2(uxxx + uux) −
t4(6uxxxxx + 20uxuxx +
10 uuxxx + 5u2ux)
(238)
(here v is defined by vx = u).
If u is subjected to zero boundary conditions u(t, − ∞) = u(t, + ∞) = 0
then KdV equation can be rewritten in Hamiltonian form
ut = {h , u}
(239)
with Poisson bivector field equal to
W = + ∞∫− ∞dx δδu∧ {δδu}x
(240)
and Hamiltonian defined by
h = + ∞∫− ∞ (ux2 − u33) dx
(241)
By taking Lie derivative of the
symplectic form along the generator of the symmetry one gets
second Poisson bivector
[E , W] =
W = + ∞∫− ∞dx ({δδu}xx ∧ {δδu}x
+ 23uδδu ∧ {δδu}x)
(242)
involved in bi-Hamiltonian structure of KdV hierarchy and
proposed by Magri [58].
Now let us show how non-Noether symmetry can be used to construct conservation laws
of KdV hierarchy. By integrating KdV it is easy to show that
J(0) = + ∞∫− ∞u dx
(243)
is conserved quantity. At the same time Lie derivative of any conserved
quantity along generator of symmetry is conserved as well,
while taking Lie derivative of J(0) along E gives rise to
infinite sequence of conservation laws J(m) = (LE)mJ(0)
that reproduce well known conservation laws of KdV equation
J(0) = + ∞∫− ∞u dx
J(1) = LEJ(0) =
¼+ ∞∫− ∞u2 dx
J(2) = (LE)2J(0) =
58+ ∞∫− ∞
(u33 − ux2) dx
J(3) = (LE)3J(0)
= 3516+ ∞∫− ∞ (536u4 −
53uux2 + uxx2) dx
J(m) = (LE)mJ(0)
(244)
Thus the conservation laws and bi-Hamiltonian structures of KdV
hierarchy are related to the non-Noether symmetry of KdV equation.
13. Nonlinear water wave equations
Among nonlinear partial differential equations that describe propagation of waves in shallow water
there are many remarkable integrable systems. We have already discussed case of KdV equation,
that possess non-Noether symmetries leading to the infinite sequence of conservation laws
and bi-Hamiltonian realization of these equations,
now let us consider other important water wave systems.
It is reasonable to start with dispersive water wave system [73],[74],
since many other models can be obtained from it by reduction.
Evolution of dispersive water wave system is governed by
the following set of equations
ut = uxw + uwx
vt = uux − vxx + 2vxw + 2vwx
wt = wxx − 2vx + 2wwx
(245)
Each symmetry of this system must satisfy linear equation
E(u)t = (wE(u))x + (uE(w))x
E(v)t = (uE(u))x − E(v)xx + 2(wE(v))x + 2(vE(w))x
E(w)t = E(w)xx − 2E(v)x + 2(wE(w))x
(246)
obtained by substituting infinitesimal transformations
u → u + zE(u) + O(z2)
v → v + zE(v) + O(z2)
w → w + zE(w) + O(z2)
(247)
into equations of motion (245) and grouping first order
(in a) terms. One of the solutions of this equation yields
the following symmetry of dispersive water wave system
E(u) = uw + x(uw)x + 2t(uw2 − 2uv + uwx)x
E(v) = 32u2 + 4vw − 3vx + x(uux + 2(vw)x − vxx)
+ 2t(u2w − uux − 3v2 + 3vw2 − 3vxw + vxx)x
E(w) = w2 + 2wx − 4v + x(2wwx + wxx − 2vx)
− 2t(u2 + 6vw − w3 − 3wwx − wxx)x
(248)
and it is remarkable that this symmetry is local in sense that E(u) in point
x depends only on u and its derivatives evaluated in the same point,
(this is not the case in KdV where symmetry is non local
due to presence of non local field v defined by vx = u).
Before we proceed let us note that dispersive water wave system is actually infinite dimensional
Hamiltonian dynamical system. Assuming that u, v and w fields
are subjected to zero boundary conditions
u(± ∞) = v(± ∞) = w(± ∞) = 0
(249)
it is easy to verify that equations (245) can be represented in Hamiltonian form
ut = {h , u}
vt = {h , v}
wt = {h , w}
(250)
with Hamiltonian equal to
h = − ¼ + ∞∫− ∞ (u2w + 2vw2 − 2vxw − 2v2)dx
(251)
and Poisson bracket defined by the following Poisson bivector field
W = + ∞∫− ∞ {½ δδu ∧ {δδu}x +
δδv ∧ {δδw}x} dx
(252)
Now using our symmetry that appears to be non-Noether, one can calculate second Poisson
bivector field involved in the bi-Hamiltonian realization of dispersive water wave system
Ŵ = [E , W] = − 2 + ∞∫− ∞
{u δδv ∧ {δδu}x
+ v δδv ∧ {δδv}x
+ {δδv}x ∧ {δδw}x
+ w δδv ∧ {δδw}x
+ {δδw}x ∧ δδw} dx
(253)
Note that Ŵ give rise to the second Hamiltonian realization of
the model
ut = {h∗ , u}∗
vt = {h∗ , v}∗
wt = {h∗ , w}∗
(254)
where
h∗ = − ¼ + ∞∫− ∞ (u2 + 2vw)dx
(255)
and { , }∗ is Poisson bracket defined by
bivector field Ŵ.
Now let us pay attention to conservation laws. By integrating third equation
of dispersive water wave system (245) it is easy to show that
J(0) =
+ ∞∫− ∞
wdx
(256)
is conservation law. Using non-Noether symmetry
one can construct other conservation laws by taking Lie derivative
of J(0) along the generator of symmetry and in this way
entire infinite sequence of conservation laws of dispersive water wave system
can be reproduced
J(0) = + ∞∫− ∞ wdx
J(1) = LEJ(0) =
− 2 + ∞∫− ∞ vdx
J(2) = LEJ(1) = (LE)2J(0) =
− 2+ ∞∫− ∞ (u2 + 2vw)dx
J(3) = LEJ(2) = (LE)3J(0) =
− 6+ ∞∫− ∞ (u2w + 2vw2 − 2vxw − 2v2)dx
J(4) = LEJ(3) = (LE)4J(0)
= − 24 + ∞∫− ∞ (u2w2 + u2wx − 2u2v − 6v2w +
2vw3 − 3vxw2 − 2vxwx)dx
J(n) = LEJ(n − 1) = (LE)nJ(0)
(257)
Thus conservation laws and bi-Hamiltonian structure of dispersive water
wave system can be constructed by means of non-Noether symmetry.
Note that symmetry (248) can be used in many other
partial differential equations that can be obtained by reduction from dispersive
water wave system. In particular one can use it in dispersiveless water wave system,
Broer-Kaup system, dispersiveless long wave system, Burger's equation etc.
In case of dispersiveless water waves system
ut = uxw + uwx
vt = uux + 2vxw + 2vwx
wt = − 2vx + 2wwx
(258)
symmetry (248) is reduced to
E(u) = uw + x(uw)x + 2t(uw2 − 2uv)x
E(v) = 32u2 + 4vw + x(uux + 2(vw)x)
+ 2t(u2w − 3v2 + 3vw2)x
E(w) = w2 − 4v + x(2wwx − 2vx) − 2t(u2 + 6vw − w3)x
(259)
and corresponding conservation laws (257) reduce to
J(0) = + ∞∫− ∞ wdx
J(1) = LEJ(0) =
− 2 + ∞∫− ∞ vdx
J(2) = LEJ(1) = (LE)2J(0) =
− 2 + ∞∫− ∞ (u2 + 2vw)dx
J(3) = LEJ(2) = (LE)3J(0) =
− 6 + ∞∫− ∞ (u2w + 2vw2 − 2v2)dx
J(4) = LEJ(3) = (LE)4J(0) =
− 24 + ∞∫− ∞ (u2w2 − 2u2v − 6v2w + 2vw3)dx
J(n) = LEJ(n − 1) = (LE)nJ(0)
(260)
Another important integrable model that can be obtained from dispersive water wave system
is Broer-Kaup system [73],[74]
vt = ½ vxx + vxw + vwx
wt = − ½ wxx + vx + wwx
(261)
One can check that symmetry (248) of dispersive water wave system,
after reduction, reproduces non-Noether symmetry of Broer-Kaup model
E(v) = 4vw + 3vx + x(2(vw)x + vxx)
+ t(3v2 + 3vw2 + 3vxw + vxx)x
E(w) = w2 − 2wx + 4v + x(2wwx − wxx + 2vx)
+ t(6vw + w3 − 3wwx + wxx)x
(262)
and gives rise to the infinite sequence of conservation laws of Broer-Kaup hierarchy
J(0) = + ∞∫− ∞ wdx
J(1) = LEJ(0) =
2 + ∞∫− ∞ vdx
J(2) = LEJ(1) = (LE)2J(0) =
4 + ∞∫− ∞ vwdx
J(3) = LEJ(2) = (LE)3J(0) =
12 + ∞∫− ∞ (vw2 + vxw + v2)dx
J(4) = LEJ(3) = (LE)4J(0) =
24 + ∞∫− ∞ (6v2w + 2vw3 + 3vxw2 − 2vxwx)dx
J(n) = LEJ(n − 1) = (LE)nJ(0)
(263)
And exactly like in the dispersive water wave system one can rewrite equations of motion
(261) in Hamiltonian form
vt = {h , v}
wt = {h , w}
(264)
where Hamiltonian is
h = ½ + ∞∫− ∞ (vw2 + vxw + v2)dx
(265)
while Poisson bracket is defined by the Poisson bivector field
W = + ∞∫− ∞ {δδv ∧ {δδw}x} dx
(266)
And again, using symmetry (262) one can recover second Poisson
bivector field involved in the bi-Hamiltonian realization of Broer-Kaup system
by taking Lie derivative of (266)
Ŵ = [E , W] = − 2 + ∞∫− ∞ {v δδv ∧ {δδv}x
− {δδv}x ∧ {δδw}x
+ w δδv ∧ {δδw}x
+ δδw ∧ {δδw}x} dx
(267)
This bivector field give rise to the second Hamiltonian realization of
the Broer-Kaup system
vt = {h∗ , v}∗
wt = {h∗ , w}∗
(268)
with
h∗ = −
¼ + ∞∫− ∞
vwdx
(269)
So the non-Noether symmetry of Broer-Kaup system yields infinite sequence
of conservation laws of Broer-Kaup hierarchy and endows it with bi-Hamiltonian structure.
By suppressing dispersive terms in Broer-Kaup system one reduces it to more simple
integarble model — dispersiveless long wave system [73],[74]
vt = vxw + vwx
wt = vx + wwx
(270)
in this case symmetry (248) reduces to more simple non-Noether symmetry
E(v) = 4vw + 2x(vw)x + 3t(v2 + vw2)x
E(w) = w2 + 4v + 2x(wwx + vx) + t(6vw + w3)x
(271)
while the conservation laws of Broer-Kaup hierarchy reduce to
sequence of conservation laws of dispersiveless long wave system
J(0) = + ∞∫− ∞ wdx
J(1) = LEJ(0) =
2 + ∞∫− ∞ vdx
J(2) = LEJ(1) = (LE)2J(0) =
4 + ∞∫− ∞ vwdx
J(3) = LEJ(2) = (LE)3J(0) =
12 + ∞∫− ∞ (vw2 + v2)dx
J(4) = LEJ(3) = (LE)4J(0) =
48 + ∞∫− ∞ (3v2w + vw3)dx
J(n) = LEJ(n − 1) = (LE)nJ(0)
(272)
At the same time bi-Hamitonian structure of Broer-Kaup hierarchy, after reduction
gives rise to bi-Hamiltonian structure of dispersiveless long wave system
W = + ∞∫− ∞ {δδv ∧ {δδw}x} dx
Ŵ = [E , W] = − 2 + ∞∫− ∞ {v δδv ∧ {δδv}x
+ w δδv ∧ {δδw}x
+ δδw ∧ {δδw}x} dx
(273)
Among other reductions of dispersive water wave system one should probably mention
Burger's equation [73],[74]
wt = wxx + wwx
(274)
However Hamiltonian realization of this equation is unknown
(for instance Poisson bivector field of dispersive water wave system
(252) vanishes during reduction).
14. Benney system
Now let us consider another integrable system of nonlinear partial
differential equations — Benney system [73],[74]. Time evolution of this dynamical
system is governed by equations of motion
ut = vvx + 2(uw)x
vt = 2ux + (vw)x
wt = 2vx + 2wwx
(275)
To determine symmetries of the system one has to look for solutions of
linear equation
E(u)t = (vE(v))x + 2(uE(w))x + 2(wE(u))x
E(v)t = 2E(u)x + (vE(w))x + (wE(v))x
E(w)t = 2E(v)x + 2(wE(w))x
(276)
obtained by substituting infinitesimal transformations
u → u + zE(u) + O(z2)
v → v + zE(v) + O(z2)
w → w + zE(w) + O(z2)
(277)
into equations (275) and grouping first order terms.
In particular one can check that the vector field E defined by
E(u) = 5uw + 2v2 + x(2(uw)x + vvx) + 2t(4uv + v2w + 3uw2)x
E(v) = vw + 6u + x((vw)x + 2ux) + 2t(4uw + 3v2 + vw2)x
E(w) = w2 + 4v + 2x(wwx + vx) + 2t(w3 + 4vw + 4u)x
(278)
satisfies equation (276) and therefore generates symmetry of Benney system.
The fact that this symmetry is local simplifies further calculations.
At the same time, it is known fact, that under zero boundary conditions
u(± ∞) = v(± ∞) = w(± ∞) = 0
(279)
Benney equations can be rewritten in Hamiltonian form
ut = {h , u}
vt = {h , v}
wt = {h , w}
(280)
with Hamiltonian
h = − ½ + ∞∫− ∞ (2uw2 + 4uv + v2w)dx
(281)
and Poisson bracket defined by the following Poisson bivector field
W = + ∞∫− ∞
{½ δδv ∧ {δδv}x
+ δδu ∧ {δδw}x} dx
(282)
Using symmetry (278) that in fact is non-Noether one, we can reproduce
second Poisson bivector field involved in the bi-Hamiltonian structure of Benney hierarchy
(by taking Lie derivative of W along E)
Ŵ = [E , W] = − 3 + ∞∫− ∞
u δδu ∧ {δδu}x
+ v δδu ∧ {δδv}x
+ w δδu ∧ {δδw}x
+ 2 δδv ∧ {δδw}x} dx
(283)
Poisson bracket defined by bivector field Ŵ gives rise
to the second Hamiltonian realization of Benney system
ut = {h∗ , u}∗
vt = {h∗ , v}∗
wt = {h∗ , w}∗
(284)
with new Hamiltonian
h∗ = 16 + ∞∫− ∞ (v2 + 2uw)dx
(285)
Thus symmetry (278) is closely related to
bi-Hamiltonian realization of Benney hierarchy.
The same symmetry yields infinite sequence of conservation laws of Benney system.
Namely one can construct sequence of integrals of motion by applying non-Noether
symmetry (278) to
J(0) = + ∞∫− ∞
wdx
(286)
(the fact that J(0) is conserved can be verified by integrating
third equation of Benney system). The sequence looks like
J(0) = + ∞∫− ∞ wdx
J(1) = LEJ(0) =
2 + ∞∫− ∞ vdx
J(2) = LEJ(1) = (LE)2J(0) =
8 + ∞∫− ∞ udx
J(3) = LEJ(2) = (LE)3J(0) =
12 + ∞∫− ∞ (v2 + 2uw)dx
J(4) = LEJ(3) = (LE)4J(0) =
48 + ∞∫− ∞ (2uw2 + 4uv + v2w)dx
J(5) = LEJ(4) = (LE)5J(0) =
240 + ∞∫− ∞ (4u2 + 8uvw + 2uw3 + 2v3 + v2w2)dx
J(n) = LEJ(n − 1) = (LE)nJ(0)
(287)
So conservation laws and bi-Hamiltonian structure of Benney hierarchy
are closely related to its symmetry, that can play important role in analysis of
Benney system and other models that can be obtained from it by reduction.
15. Conclusions
The fact that many important integrable models, such as Korteweg-de Vries equation,
Broer-Kaup system, Benney system and Toda chain,
possess non-Noether symmetries that can be effectively used
in analysis of these models, inclines us to think that non-Noether symmetries can play
essential role in theory of integrable systems and properties of this class of symmetries
should be investigated further.
The present review indicates that in many cases non-Noether symmetries lead to maximal involutive
families of functionally independent conserved quantities and in this way ensure integrability
of dynamical system. To determine involutivity of conservation laws in cases when it can not be checked
by direct computations (for instance one can not check directly the involutivity
in many generic n-dimensional models like Toda chain
and infinite dimensional models like KdV hierarchy)
we propose analog of Yang-Baxter equation, that being satisfied by
generator of symmetry, ensures involutivity of family
of conserved quantities associated with this symmetry.
Another important feature of non-Noether symmetries is their relationship with
several essential geometric concepts, emerging in theory of integrable systems, such as
Frölicher-Nijenhuis operators, Lax pairs, bi-Hamiltonian structures and
bicomplexes. On the one hand this relationship enlarges possible scope of
applications of non-Noether symmetries in Hamiltonian dynamics and on the other hand it
indicates that existence of invariant Frölicher-Nijenhuis operators,
bi-Hamiltonian structures and bicomplexes in many cases can be considered as manifestation
of hidden symmetries of dynamical system.
Acknowledgements. Author is grateful to George Jorjadze, Zakaria Giunashvili and
Michael Maziashvili for constructive discussions and help.
This work was supported by INTAS (00-00561).
[1]
M. Antonowicz, A. Fordy,
Coupled KdV equations with multi-Hamiltonian structures,
Physica D, 28 345-357,
1987
[2]
V. Arnold,
Dynamical systems III, 2nd ed.,
EMS, Springer-Verlag,
1993
[3]
V. Arnold, S. Novikov,
Dynamical systems IV, 2nd ed.,
EMS, Springer-Verlag,
2001
[4]
G. Bluman, S. Kumei,
Symmetries and differential equations,
Springer-Verlag, New York,
1989
[5]
A. Bocharov et al.,
Symmetries and conservation laws for differential equations of mathematical physics,
AMS, TMM v. 182,
1999
[6]
R. Brouzet,
Sur quelques propriétés géométriques des variétés bihamiltoniennes,
C. R. Acad. Sci. Paris 308, série I, 287-92,
1989
[7]
R. Brouzet,
Systèmes biHamiltonienes et complète intégrabilité en dimension 4,
Comptes Rendus Acad. Sci. Paris 311 895-898,
1990
[8]
J. Cariñena, L. Ibort,
Non-Noether constants of motion,
J. Phys. A:Math. Gen. 16, no. 1, 1-7,
1983
[9]
J. Cariñena, G. Marmo, M. Rañada,
Non-symplectic symmetries and bi-Hamiltonian structures of rational harmonic oscillator,
J. of Phys. A: Math. Gen. 35 L679-686,
2002
[10]
P. Casati, F. Magri, M. Pedroni,
BiHamiltonian manifolds and Sato's equations,
in O. Babelon et al., Integrable systems,
1993
[11]
G. Chavchanidze,
Non-Noether symmetries in integrable models,
J. Phys. A: Math. Gen. 37 2253-2260, math-ph/0307018,
2004
[12]
G. Chavchanidze,
Non-Noether symmetries and their influence on phase space geometry,
J. Geom. Phys. 48, 190-202,
2003
[13]
G. Chavchanidze, ,
Bi-Hamiltonian structure as a shadow of non-Noether symmetry,
Georgian Math. J. 10, 057-061,
2003
[14]
G. Chavchanidze,
Non-Noether symmetries in singular dynamical systems,
Georgian Math. J. 8 (2001) 027-032,
2001
[15]
A. Chowdhury,
Lie algebraic methods in integrable systems,
Longman Group,
2000
[16]
M. Crampin,
A note on non-Noether constants of motion,
Phys. Lett. A, 95, 209-212,
1983
[17]
M. Crampin, W. Sarlet, G. Thompson,
Bi-differential calculi and bi-Hamiltonian systems,
J. of Phys. A: Math. Gen. 33, L177-180,
2000
[18]
M. Crampin, W. Sarlet, G. Thompson,
Bi-differential calculi, bi-Hamiltonian systems and conformal Killing tensors,
J. of Phys. A: Math. Gen. 33, 8755-8770,
2000
[19]
P. Damianou,
Multiple Hamiltonian structures for Toda-type systems,
J. Math. Phys. 35 (10) 5511-5541,
1994
[20]
P. Damianou,
Symmetries of Toda equations,
J. of Phys. A: Math. Gen. 26, 3791-3796,
1993
[21]
P. Damianou, C. Sophocleous,
Master and Noether symmetries for the Toda lattice,
Proceedings of the 16th International Symposium on Nonlinear Acoustics, vol 1, 618-622 ,
2003
[22]
A. Das,
Integrable models,
World Scientific Lecture Notes in Physics, Vol. 30,
1989
[23]
L. Dickey,
Soliton equations and Hamiltonian systems,
World Scientific,
2003
[24]
A. Dimakis, F. Müller-Hoissen,
Bicomplexes and integrable models,
nlin.SI/0006029,
2000
[25]
A. Dimakis, F. Müller-Hoissen,
Bi-differential calculi and integrable models,
J. of Phys. A: Math. Gen. 33 210-23,
2000
[26]
V. Drinfeld, V. Sokolov,
Lie algebras and equations of Korteweg-de Vries type,
J. Sov. Math. 30, 1975-2036,
1985
[27]
G. Falqui, F. Magri, M. Pedroni,
The method of Poisson pairs in the theory of nonlinear PDEs,
nlin.SI/0002009,
2000
[28]
G. Falqui, F. Magri, M. Pedroni,
Bihamiltonian geometry and separation of variables for Toda lattices,
J. Nonlinear Math. Phys. 8, suppl., 118-127,
2001
[29]
G. Falqui, M. Pedroni,
Separation of variables for bi-Hamiltonian systems,
Math. Phys. Anal. Geom. 6, no. 2, 139-179,
2003
[30]
R. Fernandes,
On the master symmetries and bi-Hamiltonian structure of the Toda lattice,
J. of Phys. A: Math. Gen. 26, 3793-3803,
1993
[31]
R. Fernandes,
Completely integrable biHamiltonian systems,
University of Minnesota,
1991
[32]
A. Fokas,
Generalized symmetries and constants of motion of evolution equations,
Lett. Math. Phys. 3, 1318-1325,
1979
[33]
B. Fuchssteiner, A. Fokas,
Symplectic structures, their Bäcklund transformations and heredity symmetries,
Physica 4D, 47-66,
1981
[34]
S. Ghosh, B. Talukdar, J. Shamanna,
On the bi-Hamiltonian structure of the KdV hierarchy,
Czechoslovak J. Phys. 53, no. 5, 425-432,
2003
[35]
G. Giachetta, L. Mangiarotti, G. Sardanashvily,
Bi-Hamiltonian partially integrable systems,
J. Math. Phys. 44, no. 5, 1984-1997,
2003
[36]
F. González-Gascón,
Geometric foundations of a new conservation law discovered by Hojman,
J. of Phys. A: Math. Gen. 27 L59-60,
1994
[37]
P. Guha,
A Note on Bidifferential Calculi and Bihamiltonian systems,
IHÉS preprint M/64,
2000
[38]
R. Hermann,
Topics in the geometric theory of integrable mechanical systems,
Math SCI Press,
1984
[39]
S. Hojman,
A new conservation law constructed without using either Lagrangians or Hamiltonians,
J. of Phys. A: Math. Gen. 25 L291-295,
1992
[40]
S. Hojman, H. Harleston,
Equivalent Lagrangians: multi-dimensional case,
J. Math. Phys., 22, 1414-1419,
1981
[41]
J. Hoppe,
Lectures on integrable systems,
Springer-Verlag,
1992
[42]
A. Ibort, F. Magri, G. Marmo,
Bihamiltonian structures and Stäckel separability,
J. Geom. Phys. 32, 210-228,
2000
[43]
N. Ibragimov,
Transformation groups applied to mathematical physics,
Reidel, Boston,
1985
[44]
Y. Kosmann-Schwarzbach,
Generalized symmetries of nonlinear partial differential equations,
Lett. Math. Phys. 3, 395-404,
1979
[45]
Y. Kosmann-Schwarzbach, F. Magri,
Poisson-Nijenhuis structures,
Ann. Inst. Poincaré, 53, 35-81,
1990
[46]
M. Leo, R. Leo, G. Soliani, L. Solombrino, G. Mancarella,
Symmetry properties and bi-Hamiltonian structure of the Toda lattice,
Lett. Math. Phys. 8, 267-272,
1984
[47]
P. Libermann, C-M. Marle,
Symplectic geometry and analytical mechanics,
Reidel, Dordrecht-Boston,
1987
[48]
C. López, E. Martinez, M. Rañada,
Dynamical symmetries, non-Cartan symmetries and superintegrability of the n-dimensional harmonic oscillator,
J. of Phys. A: Math. Gen. 32, 1241-1249,
1999
[49]
M. Lutzky,
Conserved quantities from non-Noether symmetries without alternative Lagrangians,
Internat. J. Non-Linear Mech. 34, no. 2, 387-390,
1999
[50]
M. Lutzky,
Conserved quantities and velocity-dependent symmetries in Lagrangian dynamics,
Internat. J. Non-Linear Mech. 33, no. 2, 393-396,
1998
[51]
M. Lutzky,
New derivation of a conserved quantity for Lagrangian systems,
J. of Phys. A: Math. Gen. 15, L721-722,
1998
[52]
M. Lutzky,
Remarks on a recent theorem about conserved quantities,
J. of Phys. A: Math. Gen. 28, L637-638,
1995
[53]
M. Lutzky,
Noncannonical symmetries, bi-Hamiltonian systems and complete integrablity,
J. Phys. Soc. Japan 54, 64-68,
1985
[54]
M. Lutzky,
New classes of conserved quantities associated with non-Noether symmetries,
J. Phys. A 15, no. 3, L87-L91,
1982
[55]
M. Lutzky,
Isospectral representations of dynamical systems,
Lett. Nuovo Cim. 35, 393-397,
1982
[56]
M. Lutzky,
Noncannonical symmetries and isospectral representations of Hamiltonian systems,
Phys. Lett. 87A, 274-276,
1982
[57]
M. Lutzky,
Origin of non-Noether invariants,
Phys. Lett. A 75, no. 1-2, 8-10,
1980
[58]
F. Magri,
A simple model of the integrable Hamiltonian equation,
J. Math. Phys. 19 no.5, 1156-1162,
1978
[59]
C-M. Marle, J. Nunes da Costa,
Master symmetries and bi-Hamiltonian structures for the relativistic Toda lattice,
J. of Phys. A: Math. Gen. 30, 7551-7556,
1997
[60]
J. Marsden, T. Ratiu,
Introduction to mechanics and symmetry,
Springer-Verlag,
1999
[61]
C. Morosi, G. Tondo,
Quasi-bihamiltonian systems and separability,
J. of Phys. A: Math. Gen. 30, 2799-2806,
1997
[62]
W. Oevel,
Topics in soliton theory and exactly solvable non-linear equations,
World Scientific,
1987
[63]
W. Oevel, B. Fuchssteiner, H. Zhang, O. Ragnisco,
Mastersymmetries, angle variables, and recursion operator of the relativistic Toda lattice,
J. Math. Phys. 30 2664-2670 ,
1989
[64]
P. Olver,
Cannonical forms of biHamiltonian systems,
in O. Babelon et al., Integrable systems,
1993
[65]
P. Olver,
Cannonical forms and integrability of biHamiltonian systems,
Phys. Lett. 148 A, 177-187,
1990
[66]
P. Olver,
Applications of Lie groups to differential equations,
GTM 107, Springer Verlag, New York,
1986
[67]
M. Rañada,
Dynamical symmetries, bi-Hamiltonian structures and superintegrable n = 2 systems,
J. Math. Phys. 41 (4) 2121-2134,
2000
[68]
A. Razumov, M. Saveliev,
Lie algebras, geometry and Toda-type systems,
Cambridge University Press,
1997
[69]
C. Rogers, W. Schief,
Bäcklund and Darboux transformations,
Cambridge University Press,
2002
[70]
A. Sergyeyev,
A simple way of making a Hamiltonian system into a bi-Hamiltonian one,
Acta Appl. Math. 83, 183-197,
2004
[71]
Y. Suris,
On the bi-Hamiltonian structure of Toda and relativistic Toda lattices,
Phys. Lett. A 180 419 ,
1993
[72]
F. Turiel,
Classification locale d'un couple de formes symplectiques Poisson-compatibles,
Comptes Rendus Acad. Sci. Paris 308, 575-578,
1989
[73]
J.P. Wang,
A list of 1 + 1 dimensional integrable equations and their properties,
J. Nonlinear Math. Phys. 9, suppl. 1, 213–233,
2002
[74]
J.P. Wang,
Symmetries and conservation laws of evolution equations,
Ph.D. Thesis, Vrije Universiteit van Amsterdam,
1998
[75]
A. Weinstein,
The local structure of Poisson manifolds,
J. Diff. Geom. 18, 523-557,
1983