Partial wave expansion

5.4 Diabatic formalism

5.4.1 Partial wave expansion

As we did in the adiabatic case, we can write the diabatic nuclear motion Schr¨odinger Eq. (5.85) as

−~² 2µ

I∇²_Rλ+ 2W^(1)d(qλ)·∇_R

λ +

ε^d(qλ)−EI

χ^d(Rλ) =0 (5.88) The modified diabatic energy matrix ε^d(qλ) that appears in Eq. (5.88) is given by

ε^d(qλ) =ε^d(qλ)− ~²

2µW^(2)d(qλ) =



 ε^d₁₁(qλ) ε^d₁₂(qλ) ε^d₂₁(qλ) ε^d₂₂(qλ)



 (5.89)

where

ε^d_ij(qλ) =ε^d_ij(qλ)− ~²

2µW^(2)d_ij (qλ) i, j = 1,2 (5.90) The two diabatic nuclear wave functions χ^d₁(Rλ) and χ^d₂(Rλ) can be expressed as linear combinations of auxiliary nuclear wave functionsχ^d,JM₁ ^ΠΓ(R_λ) andχ^d,JM₂ ^ΠΓ(R_λ) respectively similar to their adiabatic counterparts (the linear combinations referred to as partial wave expansions and the individual χ^d,JM₁ ^ΠΓ and χ^d,JM₂ ^ΠΓ referred to as partial waves),



 χ^d₁(R_λ) χ^d₂(Rλ)



= X∞ J=0

MX=J M=−J

C_λ^JM X1 Π=0

X

Γ



 χ^d,JMΠΓ₁ (R_λ) χ^d,JMΠΓ₂ (Rλ)



 (5.91)

If χ^d,JMΠΓ₁ (Rλ) and χ^d,JM₂ ^ΠΓ(Rλ) are defined as components of the column vector χ^d,JMΠΓ(Rλ), then we can write Eq. (5.91) as

χ^d(Rλ) = X∞ J=0

MX=J M=−J

C_λ^JM X1 Π=0

X

Γ

χ^d,JMΠΓ(Rλ) (5.92)

If we define another nuclear wave function column vector

χ^d,JMΠΓ(Rλ) =



 χ^d,JMΠΓ₁ (Rλ) χ^d,JMΠΓ₂ (Rλ)



 (5.93)

thenχ^d,JMΠΓis a simultaneous eigenfunction of the diabatic matrixHb^d(R_λ) expressed as

Hb^d(R_λ) =

−~² 2µ

I∇²_Rλ+ 2W^(1)d(q_λ)·∇_R

λ +ε^d(q_λ)

, (5.94)

of the square of the total nuclear orbital angular momentum ˆJ, of its space-fixed z-component ˆJz and of the inversion operator ˆI of the nuclei through their center of mass according to the expressions

Hb^nuχ^d,JMΠΓ = Eχ^d,JMΠΓ

Jˆ²χ^d,JMΠΓ = J(J+ 1)~²χ^d,JMΠΓ Jˆ_zχ^d,JMΠΓ = M~χ^d,JMΠΓ

Iχˆ ^d,JMΠΓ = (−1)^Πχ^d,JMΠΓ

(5.95)

In these equations, J, M,Π,Γ carry the same definitions as they did in the adiabatic language. The irreducible representation Γ again refers to the fact thatχ^d,JMΠΓ leads to an electronuclear wave function which transforms according to that representation of the system’s permutation group.

Let us now expand the two nuclear motion partial waves χ^d,JM₁ ^ΠΓ and χ^d,JM₂ ^ΠΓ according to the following vector equation



 χ^d,JM₁ ^{ΠΓ,n0λΩ0λ}(ρ, ξ_λ) χ^d,JM₂ ^{ΠΓ,n0λΩ0λ}(ρ, ξλ)



=ρ^−5/2X

Ωλ

D^J_MΩ^Π_λ(ξ_λ⁽¹⁾)



 P

n1λb^d,JΠΓn_1,n1 ^0λΩ0λ

λ,Ωλ (ρ; ¯ρ)Φ^d,ΠΓ_1,n1

λ,Ωλ(ξ⁽²⁾_λ ; ¯ρ) P

n2λb^d,JΠΓn_2,n2 ^0λΩ0λ

λ,Ωλ (ρ; ¯ρ)Φ^d,ΠΓ_2,n2

λ,Ωλ(ξ⁽²⁾_λ ; ¯ρ)



 (5.96) In Eq. (5.96), Φ^d,ΠΓ_1,n1

λ,Ωλ(ξ⁽²⁾_λ ; ¯ρ) and Φ^d,ΠΓ_2,n2

λ,Ωλ(ξ_λ⁽²⁾; ¯ρ) are the diabatic two-dimensional local hyperspherical surface functions (LHSFs) that depend parametrically on ¯ρ and are defined as the eigenfunctions of a diabatic reference Hamiltonian matrix ˆh^Ω_d^λ(ξ_λ⁽²⁾; ¯ρ).

Similar to the adiabatic langiage, this matrix can be chosen to be block diagonal, i.e.,

hˆ^Ω_d^λ(ξ_λ⁽²⁾; ¯ρ) = Λˆ²₁(ξ_λ⁽²⁾) 2µρ¯²



 1 0 0 1



+



 ε^d₁₁(¯ρ, ξ_λ⁽²⁾) 0 0 ε^d₂₂(¯ρ, ξ_λ⁽²⁾)



 (5.97)

or have the off-diagonal nonadiabatic couplings built in, i.e.,

hˆ^Ω_d^λ(ξ_λ⁽²⁾; ¯ρ) = Λˆ²₁(ξ_λ⁽²⁾) 2µρ¯²



 1 0 0 1



+



 ε^d₁₁(¯ρ, ξ_λ⁽²⁾) ε^d₁₂(¯ρ, ξ_λ⁽²⁾) ε^d₂₁(¯ρ, ξ_λ⁽²⁾) ε^d₂₂(¯ρ, ξ_λ⁽²⁾)



 (5.98)

In Eqs. (5.97) and (5.98), ˆΛ²(ξ_λ⁽²⁾) refers to angular operators in θ, φ_λ coordinates [Eq. (5.44)] in the strong interaction region or ωλ, γλ coordinates [Eq. (5.46)] in the weak interaction and asymptotic regions [18, 72].

For the ˆh^Ω_d^λ(ξ_λ⁽²⁾; ¯ρ) matrix in Eq. (5.97), Φ^d,ΠΓ_1,n1

λ,Ωλ(ξ_λ⁽²⁾; ¯ρ) and Φ^d,ΠΓ_2,n2

λ,Ωλ(ξ_λ⁽²⁾; ¯ρ) are solutions of uncoupled second-order partial differential equations, whereas for the hˆ^Ω_d^λ(ξ_λ⁽²⁾; ¯ρ) matrix in Eq. (5.98), they are solutions of coupled differential equations.

Like their adiabatic partners, their calculation requires a larger computational effort than to obtain the former and as the off-diagonal couplings are built into Eq. (5.98), a smaller number of the corresponding LHSFs will be needed for convergence of the solutions of the scattering equations, as opposed to the ones resulting from Eq. (5.97), which don’t have these terms built in. Here also it is desirable to use LHSFs obtained from Eq. (5.98) rather than Eq. (5.97).

With either of these diabatic reference Hamiltonians, the LHSFs satisfy the eigen- value equation

hˆ^Ω_d^λ(ξ_λ⁽²⁾; ¯ρ)



 Φ^d,ΠΓ_1,n1

λ,Ωλ(ξ_λ⁽²⁾; ¯ρ) Φ^d,ΠΓ_2,n2

λ,Ωλ(ξ_λ⁽²⁾; ¯ρ)



=



 ^d,ΠΓ_1,n1

λ,Ωλ(¯ρ)Φ^d,ΠΓ_1,n1

λ,Ωλ(ξ_λ⁽²⁾; ¯ρ) ^d,ΠΓ_2,n2

λ,Ωλ(¯ρ)Φ^d,ΠΓ_2,n2

λ,Ωλ(ξ_λ⁽²⁾; ¯ρ)



 (5.99)

The diabatic LHSFs are not allowed to diverge anywhere on the half-sphere of fixed radius ¯ρ. This boundary condition furnishes the quantum numbers n1λ andn2λ, each of which is two-dimensional since the reference Hamiltonian ˆh^Ω_d^λ has two angular

degrees of freedom. The superscripts n⁰_λ,Ω⁰_λ in Eq. (5.96), with n⁰_λ referring to the union of n⁰_1λ and n⁰_2λ, indicate that the number of linearly independent solutions of Eqs. (5.95) is equal to the number of diabatic LHSFs used in the expansions of Eq. (5.96).

In the strong interaction region, the diabatic functions Φ^d,ΠΓ_i,n

iλ,Ωλ(θ, φλ; ¯ρ), (i = 1,2), which are eigenfunctions of the reference Hamiltonian given by Eqs. (5.97) or (5.98) and Eq. (5.44), are obtained using the same body-fixed basis Φ^d,ΠΓ_i,niλ,Ωλ(θ, φλ) used in the adiabatic representation,

Φ^d,ΠΓ_i,niλ,Ωλ(θ, φλ) = f_n^Ω_iθλ^λ (θ) g^d,ΠΓ_i,n

iφλ,Ωλ(φλ) (5.100) where niλ ≡ (niθλ, niφλ). g_i,n^d,ΠΓ

iφλ,Ωλ are again chosen to be simple trigonometric func- tions ofφλ andf_n^Ω_iθλ^λ are chosen as Jacobi polynomials in cosθ, such that the resulting diabatic nuclear wave functions transform under the operations of the permutation symmetry group of identical atoms in the way described after Eqs. (5.95). Eqs. (5.99) are then transformed into an algebraic eigenvalue eigenvector equation involving the coefficients of these expansions, which is solved numerically by linear algebra meth- ods. As was the case in the adiabatic treatment, the singularities in the ˆΛ²₁operator of Eq. (5.97) or Eq. (5.98) atθ = 0 andθ =π/2 need to be handled carefully. With the reference Hamiltonian mentioned above, the Jacobi polynomials are unable to handle the pole atθ = 0. Besides, in the current diabatic case, the second-derivative coupling element, that has been added to the diabatic energies [see Eq. (5.89)], is not singular at θ = 0. This might require using the full ˆΛ² operator and related hyperspher- ical harmonics [85, 86]. Appendix 5.A describes a method to obtain trigonometric functions in φλ that provide diabatic electronuclear functions with P3 permutation symmetries, in the presence and absence of the conical intersection.

In the weak interaction region, where the coordinates (ρ, ωλ, γλ) of Eq. (5.5) are used, the diabatic LHSFs Φ^d,ΠΓ_i,niλ,Ωλ(ωλ, γλ; ¯ρ), (i = 1,2) are eigenfunctions of the reference Hamiltonian given by Eqs. (5.97) or (5.98) and Eq. (5.46). These weak in-

teraction region diabatic LHSFs are expanded in a body-fixed basis Φ^d,ΠΓ_i,niλ,Ωλ(ωλ, γλ), used in the adiabatic language and constructed from the direct product of the asso- ciated Legendre functions of cosγλ and at a set of functions of ωλ determined by the numerical solution of a one-dimensional eigenfunction equation in ωλ using Gegen- bauer polynomials [87] in cosωλ. As in the adiabatic case, the ωλ = 0 singularity in Λˆ²₁ operator is compensated by the well-behaved Gegenbauer polynomials.

Once the diabatic LHSFs are known, they provide the basis of functions in terms of which the expansion in Eq. (5.96) is defined. The diabatic nuclear wave function vector of that equation is then inserted into the first equation of Eqs. (5.95). Use of the orthonormality of the symmetrized Wigner functions [Eq. (5.41)] and integration over the two-dimensional diabatic LHSFs, yields a set of coupled hyperradial second- order ordinary differential equations (also called coupled-channel equations) in the coefficientsb^d,J_1,n1^{ΠΓn0λΩ0λ}

λ,Ωλ (ρ; ¯ρ) andb^d,JΠΓn_2,n2 ^0λΩ0λ

λ,Ωλ (ρ; ¯ρ) analogous to the adiabatic Eq. (5.50).

Let us define the diabatic column vectors b^d,JΠΓn_i ^0λΩ0λ(ρ; ¯ρ) (i = 1,2) as the vectors whose elements are scanned by niλ,Ωλ considered as a single row index. Let us also define a matrix B^d,JΠΓ(ρ; ¯ρ) whose n⁰_λ,Ω⁰_λ column vector is obtained by stacking the vector b^d,JΠΓn₂ ^0λΩ0λ(ρ; ¯ρ) under the vector b^d,JΠΓn₁ ^0λΩ0λ(ρ; ¯ρ). These vectors, for different n⁰_λ,Ω⁰_λ, are then placed side by side thereby generating a square matrix B^d,JΠΓ whose dimensions are the total number of LHSFs (channels) used. The coupled hyperradial equation satisfied by this matrix has the form

−~² 2µI d²

dρ² +V^d,JΠΓ(ρ; ¯ρ)

B^d,JΠΓ(ρ; ¯ρ) =EB^d,JΠΓ(ρ; ¯ρ) (5.101) where V^d,JΠΓ(ρ; ¯ρ) is the interaction potential matrix obtained by this derivation procedure and which encompasses ^d(¯ρ):

V^d,JΠΓ(ρ; ¯ρ) =



 V^d,JΠΓ₁₁ (ρ; ¯ρ) V₁₂^d,JΠΓ(ρ; ¯ρ) V^d,JΠΓ₂₁ (ρ; ¯ρ) V₂₂^d,JΠΓ(ρ; ¯ρ)



 (5.102)

Its dimensions are those of B^d,JΠΓ(ρ; ¯ρ).