Free particle on SU(2) group manifold George Chavchanidze Department of Theoretical Physics A. Razmadze Institute of Mathematics 1 Aleksidze Street Tbilisi 0193 Georgia Abstract

In the present paper classical and quantum dynamics of a free particle on SU(2) group manifold is considered. Poisson structure of the dynamical system and commutation relations for generalized momenta are derived. Quantization is carried out and the eigenfunctions of the Hamiltonian are constructed in terms of coordinate free objects. SU(2)/U(1) coset model yielding after Hamiltonian reduction free particle on S2 sphere is considered and Hamiltonian reduction of coset model is carried out on both classical and quantum level.

Keywords: Dynamics on group manifold, Quantization on group manifold,

MSC 2000: 70H33; 70H06; 53Z05

Lagrangian description

The dynamics of a free particle on SU(2) group manifold is described by the Lagrangian L = 〈g− 1ġg− 1ġ〉 where g ∈ SU(2) and 〈 〉 denotes the normalized trace 〈 · 〉 = − ½Tr( · ) which defines a scalar product in su(2) algebra. This Lagrangian gives rise to equations of motion ddtg− 1ġ = 0 that describe dynamics of particle on group manifold. Also, one can notice that it has SU(2) "right" and SU(2) "left" symmetry. It means that it is invariant under the following transformations g       →      h1g g       →    gh2 where h1, h2 ∈ SU(2)

According to the Noether's theorem these symmetries lead to the matrix valued conserved quantities C = g− 1ġ           ddtC = 0 and S = ġg− 1           ddtS = 0 To construct integrals of motion out of C and S let us introduce the basis of su(2) algebra — three matrices: T1 = i 0 0 − i       T2 = 0 − 1 1 0       T3 = 0 i i 0 The elements of su(2) are traceless anti-hermitian matrices, and any A ∈ su(2) can be parameterized in the following way A = AnTn            n = 1, 2, 3 Scalar product AB = 〈AB〉 = − ½Tr(AB) ensures that An = 〈ATn〉            (〈TnTm〉 = δnm) Now we can introduce six functions Cn = 〈TnC〉          n = 1, 2, 3           C = CnTn Sn = 〈TnS〉           n = 1, 2, 3            S = SnTn which are integrals of motion.

Conservation of C and S leads to general solution of Euler-Lagrange equations ddtg− 1ġ = 0           ⇒           g− 1ġ = const g = eCtg(0) These are well known geodesics on Lie group.

Hamiltonian description

Working in a first order Hamiltonian formalism we can construct new Lagrangian which is equivalent to the initial one Λ = 〈C(g− 1ġ − v)〉 + ½〈v2 in sense that variation of C provides g− 1ġ = v and Λ reduces to L. Variation of v gives C = v and therefore we can rewrite equivalent Lagrangian Λ in terms of C and g variables Λ = 〈Cg− 1ġ〉 − ½ 〈C2 where function H = ½〈C2 plays the role of Hamiltonian and one-form 〈Cg− 1dg〉 is a symplectic potential θ. External differential of θ is the symplectic form ω = dθ = − 〈g− 1dg ∧ dC〉 − 〈Cg− 1 dg ∧ g− 1dg〉 that determines Poisson brackets, the form of Hamilton's equation and provides isomorphism between vector fields and one-forms X      →      iXω For any smooth SU(2) valued smooth function f ∈ SU(2) one can define Hamiltonian vector field Xf by iXfω = − df where iXω denotes the contraction of X with ω. According to its definition Poisson bracket of two functions is {f , g} = LXfg = iXfdg = ω(Xf , Xg) where LXfg denotes Lie derivative of g with respect to vector filed Xf. The skew symmetry of ω provides skew symmetry of Poisson bracket.

Hamiltonian vector fields that correspond to Cn, Sm and g functions are Xn = XCn = ([C ,Tn] , gTn) Ym = XSm = ([C , gTmg− 1] , Tmg ) and give rise to the following commutation relations {Sn , Sm} = − 2εnmk Sk {Cn , Cm} = 2εnmk Ck {Cn , Sm} = 0 {Cn , g} = gTn {Sm , g} = Tmg The results are natural. C and S that correspond respectively to the "right" and "left" symmetry commute with each other and independently form su(2) algebras. Now knowing Poisson bracket structure one can write down Hamilton's equations ġ = {H , g} = gR Ċ = {H , C} = 0

Quantization

Let's introduce operators Ĉn = i2LXn Ŝm = − i2LYm They act on the square integrable functions (see Appendix A) on SU(2) and satisfy quantum commutation relations n , Ŝm] = iεnmk Ŝk n , Ĉm] = iεnmk Ĉk n , Ŝm] = 0 The Hamiltonian is defined as Ĥ = Ĉ2 = Ŝ2 and the complete set of observables that commute with each other is Ĥ,           Ĉa,            Ŝb with some fixed a and b. Using a simple generalization of a well known algebraic construction (see Appendix B) one can check that the eigenvalues of the quantum observables Ĥ, Ĉa and Ŝb have the form Ĥψjsc = j(j + 1)ψjsc where j takes positive integer and half integer values j = 0, 12, 1, 32, 2 ... Ĉaψjsc = cψjsc Ŝbψjsc = sψjsc with c and s taking values in the following range − j, − j + 1, ... , j − 1, j Further we construct the corresponding eigenfunctions ψjsc. The first step of this construction is to note that the function 〈Tg〉 where T = (1 + iTa)(1 + iTb) is an eigenfunction of Ĥ, Ĉa and Ŝb with eigenvalues ¾, ½, ½ respectively. Proof of this proposition is straightforward. Using 〈Tg〉 one can construct the complete set of eigenfunctions of Ĥ, Ĉa and Ŝb operators ψjsc = Ŝj − sĈj − c〈Tg〉2j in the manner described in Appendix B.

Free particle on S² as a SU(2)/U(1) coset model

Free particle on 2D sphere can be obtained from our model by gauging U(1) symmetry. In other words let's consider the following local gauge transformations g      →      h(t)g Where h(t) ∈ U(1) ⊂ SU(2) is an element of U(1). Without loss of generality we can take h = eβ(t)T3 Since T3 is antihermitian h(t) ∈ U(1) and since h(t) depends on t Lagrangian L = 〈g− 1ġg− 1ġ〉 is not invariant under (38) local gauge transformations.

To make (40) gauge invariant we should replace time derivative with covariant derivative ddtg     →    ∇g = ddt + Bg where B can be represented as follows B = bT3 ∈ su(2) with transformation rule B      →     hBh− 1dhdth− 1 or in terms of b variable b    →     b − dt The new Lagrangian LG = 〈g− 1∇gg− 1∇g〉 is invariant under (38) local gauge transformations. But this Lagrangian as well as every gauge invariant Lagrangian is singular. It contains additional non-physical degrees of freedom. To eliminate them we should eliminate B using Lagrange equations ∂LG∂B      →      b = − 〈ġg− 1T3 put it back in (45) and rewrite last obtained Lagrangian in terms of gauge invariant variables. LG = 〈(g− 1ġ − S3T3)2 It's obvious that the following Z = g− 1T3g ∈ su(2) element of su(2) algebra is gauge invariant. Since Z ∈ su(2) it can be parameterized as follows Z = zaTa where za are real functions on SU(2) za = 〈ZTa

So we have three gauge invariant variables za (a = 1, 2, 3) but it's easy to check that only two of them are independent. Indeed 〈Z2〉 = 〈g− 1T3gg− 1T3g〉 = 〈T32〉 = 1 otherwise 〈Z2〉 = 〈zaTazbTb〉 = zaza

So configuration space of SU(2)/U(1) coset model is sphere. By direct calculations one can check that after being rewritten in terms of gauge invariant variables LG takes the form LG = ¼〈Z− 1ŻZ− 1Ż〉 This Lagrangian describes free particle on the sphere. Indeed, since Z = zaTa it's easy to show that LG = ¼〈Z− 1ŻZ− 1Ż〉 = ¼〈ZŻZŻ〉 = ½żaża So SU(2)/U(1) coset model describes free particle on S2 manifold.

Quantization of the coset model.

Working in a first order Hamiltonian formalism one can introduce equivalent Lagrangian ΛG = 〈C(g− 1ġ − u)〉 + ½ 〈(u + g− 1Bg)2 variation of u provides C = u + g− 1Bg u = C − g− 1Bg Rewriting ΛG in terms of C and g leads to ΛG = 〈Cg− 1ġ〉 − ½ 〈C2〉 − 〈BgCg− 1〉 = 〈Cg− 1ġ〉 − ½ 〈C2〉 − b〈gCg− 1T3〉 = 〈Cg− 1ġ〉 − ½ 〈C2〉 − bS3 Due to the gauge invariance of ΛG we obtain constrained Hamiltonian system, where 〈Cg− 1dg〉 is symplectic potential, function H = ½〈C2 plays the role of Hamiltonian and b is a Lagrange multiple leading to the first class constrain φ = 〈gCg− 1T3〉 = 〈ST3〉 = S3 = 0 So coset model is equivalent to the initial one with (59) constrain. Using technique of the constrained quantization, instead of quantizing coset model we can subject quantum model that corresponds to the free particle on SU(2), to the following operator constrain Ŝ3|ψ〉 = 0 Hilbert space of the initial system, that is linear span of ψjcs           j = 0, 12, 1, 32, 2, ... wave functions, reduces to the linear span of ψjc0            j = 0, 1, 2, 3, ... wave functions. Indeed, Ŝ3ψjcs = 0 implies s = 0, and if s = 0 then j is integer. Thus c takes − j, − j + 1, ..., j − 1, j integer values only. Wave functions ψjcs rewriten in terms of gauge invariant variables up to a constant multiple should coincide with well known spherical harmonics ψjc0 ∼ Jjc One can check the following ψjc0 ∼ ŜjĈj − c 〈Tg〉2j ∼ Ĉj − c 〈T+g− 1T3g〉j ∼ Ĉj − csinjθeijθ ∼ Ĉj − cJjj ∼Jjc This is an example of using large initial model in quantization of coset model.

Appendix A

Scalar product in Hilbert space is defined as follows 〈ψ12〉 = SU(2) a = 13 〈g− 1dgTa〉(ψ1)ψ2 It's easy to prove that under this scalar product operators Ĉn and Ŝm are hermitian. Indeed 〈ψ1nψ2〉 = SU(2) a = 13 〈g− 1dgTa〉(ψ1) i2LXnψ2 = SU(2) a = 13 〈g− 1dgTai2LXnψ1ψ2 Where integration by part has been used and the additional term coming from measure a = 13 〈g− 1dgTa vanished since LXn〈g− 1dgTa〉 = 0 For more transparency one can introduce the following parameterization of SU(2). For any g ∈ SU(2). g = eqaTa Then the symplectic potential takes the form 〈Cg− 1dg〉 = Cadqa and scalar product becomes 〈ψ12〉 = 0 0 0 d3q(ψ1)ψ2 that coincides with (65) because of dqa = 〈g− 1dg Ta

Appendix B

Without loss of generality we can take Ĥ, Ŝ3 and Ĉ3 as a complete set of observables. Assuming that operators Ĥ, Ŝ3 and Ĉ3 have at least one common eigenfunction Ĥψ = Eψ Ĉ3ψ = cψ Ŝ3ψ = sψ it is easy to show that eigenvalues of Ĥ are non-negative E ≥ 0 and conditions E − c2 ≥ 0 E − s2 ≥ 0 are satisfied. Indeed, operators Ĉ and Ŝ are selfadjoint so 〈ψ|Ĥ|ψ〉 = 〈ψ|Ĉ2|ψ〉 = 〈ψ|ĈaĈa|ψ〉 = 〈ψ|(Ĉa)Ĉa|ψ〉 = 〈Ĉaψ|Ĉaψ〉 = ∥Ĉaψ∥ ≥ 0 To prove (74) we shall consider Ĉ12 + Ĉ22 and Ŝ12 + Ŝ22 operators 〈ψ|Ĉ12 + Ĉ22|ψ〉 = ∥Ĉ1 ψ∥ + ∥Ĉ2 ψ∥ ≥ 0 and 〈ψ|Ĉ12 + Ĉ22|ψ〉 = 〈ψ|Ĥ − Ĉ32|ψ〉 = (E − c2)〈ψ|ψ〉 thus E − c2 ≥ 0.

Now let's introduce new operators Ĉ+ = iĈ1 + Ĉ2            Ĉ = iĈ1 − Ĉ2 Ŝ+ = iŜ1 + Ŝ2            Ŝ = iŜ1 − Ŝ2 These operators are not selfadjoint, but ) = Ĉ+ and ) = Ŝ+ and they fulfill the following commutation relations ± , Ĉ3] = ± Ĉ±           [Ŝ± , Ŝ3] = ± Ŝ± + , Ĉ] = 2Ĉ3           [Ŝ+ , Ŝ] = 2Ŝ3 , Ŝ] = 0 where takes values +, −, 3 using these commutation relations it is easy to show that if ψλcs is eigenfunction of Ĥ, Ŝ3 and Ĉ3 with corresponding eigenvalues : Ĥψλcs = λψλcs Ŝ3ψλcs = sψλcs Ĉ3ψλcs = cψλcs then Ĉ±ψλcs and Ŝ±ψλcs are the eigenfunctions with corresponding eigenvalues λ, s ± 1, c and λ , s, c ± 1. Consequently using Ĉ±, Ŝ± operators one can construct a family of eigenfunctions with eigenvalues c, c ± 1, c ± 2, c ± 3, ... s, s ± 1, s ± 2, s ± 3, ... but conditions (74) give restrictions on a possible range of eigenvalues. Namely we must have λ − c2 ≥ 0 λ − s2 ≥ 0 In other words, in order to interrupt (84) sequences we must assume Ŝ+ ψλcj = 0            Ŝψλc, − j = 0 Ĉ+ψλks = 0           Ĉψλ, − ks = 0 for some j and k, therefore s and c could take only the following values − j, − j + 1, ... , j − 1, j − k, − k + 1, ... , k − 1, k The number of values is 2j + 1 and 2k + 1 respectively. Since number of values should be integer, j and k should take integer or half integer values j = 0, 12, 1, 32, 2, ... k = 0, 12, 1, 32, 2, ... Now using commutation relations we can rewrite Ĥ in terms of Ĉ±, Ĉ3 operators Ĥ = Ĉ+ Ĉ + Ĉ32 + Ĉ3 and it is clear that j = k and λ = j(j + 1) = k(k + 1)

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