Introduction to many-body Green-function theory

Transcription

1 Introduction to many-body Green-function theory Julien Toulouse Laboratoire de Chimie Théorique Université Pierre et Marie Curie et CNRS, Paris, France July 15, 2015 Contents 1 A tour of many-body Green function theory One-particle Green function Two-particle Green function and linear-response function Dyson equation Quasiparticle equations Bethe-Salpeter equation Hartree-Fock approximation One-particle Green function Linear-response function Linear-response equation in a spin-orbital basis Spin-adapted linear-response equation for closed-shell systems Example: H 2 in a minimal basis Appendices 15 A Fourier-transform formulas B Equation of motion for the one-particle Green function C Schwinger functional derivative technique References 19 1

2 These notes provide an introduction to many-body Green function theory, with some links to the quantum chemistry perspective. For more details, see Refs. [1 7]. 1 A tour of many-body Green function theory We consider a many-body Hamiltonian operator expressed in real-space second quantization Ĥ = dxˆψ (x)h(r)ˆψ(x)+ 1 dx 1 dx 2 2 ˆψ (x 1 )ˆψ (x 2 )v(r 1,r 2 )ˆψ(x 2 )ˆψ(x 1 ), (1) whereh(r) = ( 1/2) 2 +v ne (r)istheone-electronhamiltonian, andv(r 1,r 2 ) = 1/ r 1 r 2 isthe electron-electron interaction. The creation ˆψ (x) and annihilation ˆψ(x) field operators written with space-spin coordinates x = (r,σ), obey usual anticommutation rules: {ˆψ (x 1 ), ˆψ (x 2 )} = {ˆψ(x 1 ), ˆψ(x 2 )} = 0 and {ˆψ (x 1 ), ˆψ(x 2 )} = δ(x 1 x 2 ). We consider a system of N electrons described by the Hamiltonian Ĥ with a normalized ground state denoted by N. The one-particle and two-particle Green functions of such a system are useful quantities that encapsulate information about the excited states of the system. 1.1 One-particle Green function The time-ordered one-particle Green function is defined as ig(1,2) = N T[ˆψ(1)ˆψ (2)] N = θ(t 1 t 2 ) N ˆψ(1)ˆψ (2) N θ(t 2 t 1 ) N ˆψ (2)ˆψ(1) N, (2) where the indices 1 and 2 stand for space-spin-time coordinates 1 = (x 1,t 1 ) and 2 = (x 2,t 2 ), T is the Wick time-ordering operator which orders the operators with larger times on the left, and θ is the Heaviside step function. The whole time-dependence is contained in the annihilation and creation field operators in the Heisenberg picture: ˆψ(1) = e iĥt 1 ˆψ(x1 )e iĥt 1 and ˆψ (2) = e iĥt 2 ˆψ (x 2 )e iĥt 2. In the particle-hole language, the first term of the right-hand-side of Eq. (2) describes the propagation of a particle (i.e., an added electron in the system), while the second term describes the propagation of a hole (i.e., a removed electron from the system). The one-particle Green function can be considered as a time-dependent generalization of the one-particle density matrix. Indeed, the one-particle density matrix n 1 (x 1,x 2 ) is just n 1 (x 1,x 2 ) = N ψ (x 2 )ψ(x 1 ) N = ig(x 1 t 1,x 2 t + 1 ), (3) where t + 1 means that an infinitesimal positive shift has been added to the time variable, t+ 1 = t , in order to select the second term in the right-hand side of Eq. (2). The one-particle density can also be obtained by setting x 1 = x 2 and summing over the spin coordinate n(r 1 ) = i dσ 1 G(r 1 σ 1 t 1,r 1 σ 1 t + 1 ). (4) In the absence of a time-dependent external potential, the system is invariant under time translation, therefore the Green function depends only on τ = t 1 t 2. By introducing the completeness relations for (N 1)- and (N +1)-electron states, one can get ig(x 1,x 2 ;τ) =θ(τ) N ˆψ(x 1 ) N +1,a N +1,a ˆψ (x 2 ) N e i(e N+1,a E N )τ a θ( τ) N ˆψ (x 2 ) N 1,i N 1,i ˆψ(x 1 ) N e i(e N E N 1,i )τ, i (5) 2

3 where E N is the energy of the N-electron ground state N, and E N+1,a and E N 1,i are the energies of the a-th (N + 1)-electron state N + 1,a and of the i-th (N 1)-electron state N 1, i, respectively. Exercise 1 : Derive Eq. (5) from Eq. (2). The Lehmann representation of the one-particle Green function is then obtained by Fourier transform (cf. Appendix A) G(x 1,x 2 ;ω) = a f a (x 1 )f a(x 2 ) ω E a +i0 + + i f i (x 1 )fi (x 2) ω E i i0 +, (6) where f a (x) = N ˆψ(x) N + 1,a and f i (x) = N 1,i ˆψ(x) N are the Dyson orbitals (also called Lehmann amplitudes or quasiparticle wave functions), and E a = E N+1,a E N = A a and E i = E N E N 1,i = I i are the quasiparticle energies, i.e. minus the electron affinities A a and ionization energies I i, respectively. Therefore, the one-particle Green function contains information relative to photoelectron and inverse photoelectron spectroscopies. Exercise 2 : Using the Fourier-transform formulas given in Appendix A, derive Eq. (6). 1.2 Two-particle Green function and linear-response function The time-ordered two-particle Green function is defined as i 2 G 2 (1,2;1,2 ) = N T[ˆψ(1)ˆψ(2)ˆψ (2 )ˆψ (1 )] N. (7) Depending on the time ordering, it describes the propagation of a pair of particles, of holes, or of a particle and a hole. A related quantity is the four-point linear-response function defined as iχ(1,2;1,2 ) = i 2 G 2 (1,2;1,2 ) ig(1,1 ) ig(2,2 ). (8) It describes the coupled motion of a pair of two particles, or two holes, or a particle and a hole, minus the motion of the independent ones. When the time variables are appropriately ordered, the four-point linear-response function reduces to the particle-hole linear-response function (or polarization propagator or, simply, linear-response function) iχ(x 1 t 1,x 2 t 2 ;x 1t + 1,x 2t + 2 ) = N T[ˆψ (x 1t 1 )ˆψ(x 1 t 1 )ˆψ (x 2t 2 )ˆψ(x 2 t 2 )] N N ˆψ (x 1t 1 )ˆψ(x 1 t 1 ) N N ˆψ (x 2t 2 )ˆψ(x 2 t 2 ) N, (9) where t + 1 = t and t + 2 = t It describes the propagation of a particle-hole pair, and it is the relevant quantity to consider for optical absorption or emission spectroscopy. Due to time translation invariance, the linear-response function depends only on τ = t 1 t 2. By introducing the completeness relation for N-electron states, one can get χ(x 1,x 2 ;x 1,x 2;τ) = θ(τ) n 0 N ˆψ (x 1)ˆψ(x 1 ) N,n N,n ˆψ (x 2)ˆψ(x 2 ) N e i(e N,n E N )τ +θ( τ) n 0 N ˆψ (x 2)ˆψ(x 2 ) N,n N,n ˆψ (x 1)ˆψ(x 1 ) N e i(e N,n E N )τ, (10) 3

4 where E N,n is the energy of n-th N-electron excited state N,n. The ground state N,0 = N is excluded from the sum. The Lehmann representation of the linear-response function is thus χ(x 1,x 2 ;x 1,x 2;ω) = [ N ˆψ (x 1 )ˆψ(x 1 ) N,n N,n ˆψ (x 2 )ˆψ(x 2 ) N ω ω n +i0 + n 0 N ˆψ (x 2 )ˆψ(x 2 ) N,n N,n ˆψ (x 1 )ˆψ(x 1 ) N ω +ω n i0 + ], (11) where ω n = E N,n E N are the excitation energies. Therefore, the linear-response function has poles in ω at the excitation energies ω n and at the de-excitation energies ω n. Exercise 3 : Derive Eq. (10) and Eq. (11). It is also useful to define the independent-particle (IP) four-point linear-response function and the corresponding IP linear-response function χ IP (1,2;1,2 ) = ig(1,2 )G(2,1 ), (12) χ IP (x 1,x 2 ;x 1,x 2;τ) = χ IP (x 1 t 1,x 2 t 2 ;x 1t + 1,x 2t + 2 ) = ig(x 1,x 2;τ)G(x 2,x 1; τ). (13) Its Lehmann representation is easily obtained by using Eq. (5) and taking the Fourier transform, χ IP (x 1,x 2 ;x 1,x 2;ω) = [ f i (x 1 )f a(x 1 )fa(x 2 )f i(x 2 ) ω (E i a a E i )+i0 + f i (x 2 )f a(x 2 )fa(x 1 )f ] i(x 1 ) ω +(E a E i ) i0 +, (14) where we see that it contains quasiparticle energy differences as poles in ω. The fundamental gap, E gap = I 0 A 0, corresponds to the smallest of such quasiparticle energy differences. Exercise 4 : Derive Eq. (14). Having seen their useful physical contents, we would like to calculate the one-particle Green function and the linear-response function without solving the full many-body problem. We will derive now the Dyson equation for calculating the one-particle Green function, and the Bethe-Salpeter equation for calculating the four-point linear-response function. 1.3 Dyson equation In Appendix B, it is shown that the one-particle Green function satisfies the following equation of motion [ i ] h(1) G(1,2)+i d3v(1,3)g 2 (1,3 + ;2,3 ++ ) = δ(1,2), (15) with the notations h(1) = h(r 1 ) and v(1,3) = v(r 1,r 3 )δ(t 1 t 3 ), and 3 + means that the time variable has an infinitesimal positive shift t + 3 = t , and 3 ++ means another shift t ++ 3 = 4

5 t Thus, theequationofmotionoftheone-particlegreenfunctioninvolvesthetwo-particle Green function. One can formally close this equation as [ i ] h(1) G(1,2) d3σ Hxc (1,3)G(3,2) = δ(1,2), (16) where Σ Hxc (1,3) is the Hartree-exchange-correlation self-energy which encompasses all the twoelectron effects. The self-energy is defined by d3σ Hxc (1,3)G(3,2) = i d3v(1,3)g 2 (1,3 + ;2,3 ++ ), (17) or, moreexplicitly, aftermultiplyingfromtherightbytheinverseofthegreenfunctiong 1 (2,4) and integrating over the coordinate 2, Σ Hxc (1,4) = i d2d3v(1,3)g 2 (1,3 + ;2,3 ++ )G 1 (2,4), (18) where we have used the definition of the inverse, d2g(3,2)g 1 (2,4) = δ(3,4). The non-interacting Green function G h corresponding to the one-electron Hamiltonian h has no two-electron contributions and therefore follows the equation of motion 1 [ i ] h(1) G h (1,2) = δ(1,2). (19) Writing G(1,2) = d3d4g h (1,4)G 1 h (4,3)G(3,2) in the first term of Eq. (16) and using Eq. (19), one gets d3 [ G 1 h (1,3) Σ Hxc(1,3) ] G(3,2) = δ(1,2), (20) which can be rewritten as, after multiplying from the left by d1g h (4,1) and renaming the indices, G(1,2) = G h (1,2)+ d3d4g h (1,3)Σ Hxc (3,4)G(4,2), (21) or, equivalently, after multiplying from the right Eq. (20) by G 1 (2,4), integrating over the coordinate 2, and renaming the indices, G 1 (1,2) = G 1 h (1,2) Σ Hxc(1,2). (22) Eq. (21) or (22) constitutes the Dyson equation for the one-particle Green function. Knowing the non-interacting Green function G h and an approximation for the self-energy Σ Hxc, it allows one to calculate the corresponding interacting Green function G. 1.4 Quasiparticle equations We will show now that the Dyson equation can be conveniently reformulated as eigenvalue equations for the Dyson orbitals and the quasiparticle energies. 1 Non-interacting one-particle Green functions are frequently denoted by G 0 in the literature. However, there are several possible non-interacting Green functions (Hartree, Hartree-Fock, Kohn-Sham, etc...), so we prefer the notation G h which explicitly specifies the corresponding one-electron Hamiltonian h. 5

6 Just like the one-particle Green function, the self-energy depends only the time difference τ = t 1 t 2, i.e. Σ Hxc (1,2) = Σ Hxc (x 1,x 2 ;τ). Fourier transforming the equation of motion (16) then gives [ω h(r 1 )]G(x 1,x 2 ;ω) dx 3 Σ Hxc (x 1,x 3 ;ω)g(x 3,x 2 ;ω) = δ(x 1,x 2 ), (23) where Σ Hxc (x 1,x 3 ;ω) is the Fourier transform of Σ Hxc (x 1,x 3 ;τ). Exercise 5 : Derive Eq. (23) from Eq. (16). Inserting the Lehmann representation (6) of G(x 1,x 2 ;ω) in Eq. (23), one can select the term corresponding to a given pole E k (assumed to be discrete and nondegenerate) of G(x 1,x 2 ;ω) by multiplying the equation by ω E k and taking the limit ω E k, giving [E k h(r 1 )]f k (x 1 )fk (x 2) dx 3 Σ Hxc (x 1,x 3 ;E k )f k (x 3 )fk (x 2) = 0, (24) or, after simplifying by f k (x 2), h(r 1 )f k (x 1 )+ dx 3 Σ Hxc (x 1,x 3 ;E k )f k (x 3 ) = E k f k (x 1 ). (25) Eq. (25) constitutes the quasiparticle equations giving the Dyson orbitals f k (x) and the corresponding quasiparticle energies E k (ionization energies or electron affinities). The self-energy Σ Hxc (x 1,x 3 ;E k ) plays the role of a nonlocal energy-dependent potential. The quasiparticle equations may be thought of as an extension of the Hartree-Fock or Kohn-Sham orbital equations. Because of the dependence on the energy E k, each Dyson orbital f k (x) is an eigenfunction of a different operator. Consequently, the Dyson orbitals do not form a basis of the one-electron space, but constitutes an overcomplete set. 1.5 Bethe-Salpeter equation The Bethe-Salpeter equation for the four-point linear-response function can be derived from the Dyson equation using the functional derivative technique of Schwinger. The idea is to add a time-dependent nonlocal external potential u(1,1 ) = u(x 1,x 1 ;t 1)δ(t 1 t 1 ) in the Hamiltonian, leading to generalized Green functions, and consider functional derivatives of Green functions with respect to u(1,1 ) to generate higher-particle Green functions. The potential u is then eventually set to zero at the end of the derivation. For example, the four-point linear-response function can be expressed as the functional derivative χ(1,2;1,2 ) = i δg(1,1 ) δu(2,2), (26) where we consider variations of the one-particle Green function δg(1,1 ) induced by a variation of the external potential δu(2,2). Assuming that the inverse of χ(1,2;1,2 ) exists, it can be written as 2 χ 1 (1,2;1,2 ) = i δu(1,1 ) δg(2,2), (27) 2 The inverse of a 4-point function χ(1,2;1,2 ) is defined according to d1 d2 χ(1,2;1,2 )χ 1 (2,1 ;4,3) = δ(1, 3)δ(2, 4). 6

7 where we consider now variations of the external potential δu(1,1 ) induced by a variation of the one-particle Green function δg(2,2). The Bethe-Salpeter can be derived by starting from the Dyson equation of Eq. (22), written in the presence of the external potential u, G 1 (1,1 ) = G 1 h (1,1 ) u(1,1 ) Σ Hxc (1,1 ), (28) and taking the functional derivative with respect to ig(2,2). Since G 1 h (1,1 ) is independent from u, its functional derivative vanishes, and it remains i δg 1 (1,1 ) δg(2,2) = i δu(1,1 ) δg(2,2) +iδσ Hxc(1,1 ) δg(2,2). (29) The term on the left-hand side of Eq. (29) can be directly calculated (see Exercise 6) i δg 1 (1,1 ) δg(2,2) = ig 1 (1,2 )G 1 (2,1 ) = χ 1 IP (1,2;1,2 ), (30) where χ 1 IP (1,2;1,2 ) is the inverse of the IP four-point linear-response function χ IP (1,2;1,2 ) defined in Eq. (12). The first term on the right-hand side of Eq. (29) is just the inverse of the four-point linear-response function χ 1 (1,2;1,2 ) according to Eq. (27). Finally, the last term on the right-hand side of Eq. (29) defines the Hartree-exchange-correlation Bethe-Salpeter kernel i δσ Hxc(1,1 ) δg(2,2) = Ξ Hxc (1,2;1,2 ). (31) We thus arrive at the Bethe-Salpeter equation χ 1 (1,2;1,2 ) = χ 1 IP (1,2;1,2 ) Ξ Hxc (1,2;1,2 ), (32) or, equivalently, after matrix multiplicating by χ IP from the left and by χ from the right, χ(1,2;1,2 ) = χ IP (1,2;1,2 )+ d3d4d5d6 χ IP (1,4;1,3)Ξ Hxc (3,6;4,5)χ(5,2;6,2 ). Knowing approximations for the IP four-point linear-response function χ IP and the Bethe- Salpeter kernel Ξ Hxc, it allows one to calculate the corresponding interacting four-point linearresponse function χ. (33) Exercise 6 : Derive Eq. (30). For this, start from d3 G 1 (1,3)G(3,4) = δ(1,4) and take the functional derivative with respect to G(2,2). 2 Hartree-Fock approximation 2.1 One-particle Green function The starting point is the Hartree-Fock (HF) approximate form for the two-particle Green function G HF 2 (1,2;1,2 ) = G(1,1 ) G(2,2 ) G(1,2 )G(2,1 ), (34) 7

8 with a direct (or Hartree) and an exchange (or Fock) term. Injecting Eq. (34) into Eq. (18) gives the HF self-energy Σ HF Hx(1,2) = Σ H (1,2)+Σ x (1,2), (35) where Σ H (1,2) is the Hartree contribution Σ H (1,2) = iδ(1,2) and Σ x (1,2) is the exchange contribution, d3g(3,3 + )v(1,3), (36) Σ x (1,2) = ig(1,2 + )v(1,2). (37) Exercise 7 : Derive Eqs. (36) and (37) from Eq. (18) and (34). The Hartree self-energy can be written more explicitly as Σ H (x 1 t 1,x 2 t 2 ) = δ(t 1 t 2 )δ(x 1 x 2 ) dr 3 n(r 3 )v(r 1,r 3 ) = δ(t 1 t 2 )δ(x 1 x 2 )v H (r 1 ), (38) where we have used Eq. (4) to introduce the density n(r 3 ), and v H (r 1 ) is the local Hartree potential. Similarly, the exchange self-energy can be written as Σ x (x 1 t 1,x 2 t 2 ) = δ(t 1 t 2 )n 1 (x 1,x 2 )v(r 1,r 2 ) = δ(t 1 t 2 )v HF x (x 1,x 2 ), (39) wherewehaveusedeq.(3)tointroducetheone-particledensitymatrixn 1 (x 1,x 2 ),andv HF x (x 1,x 2 ) is the nonlocal HF exchange potential. The delta function δ(t 1 t 2 ) in Eqs. (38) and (39) implies that the Fourier transforms of Σ H and Σ x are independent from the frequency ω. The quasiparticle equations (25) just reduce to the usual HF equations h(r 1 )ψ k (x 1 )+v H (r 1 )ψ k (x 1 )+ dx 3 vx HF (x 1,x 3 )ψ k (x 3 ) = ε k ψ k (x 1 ), (40) where ψ k (x 1 ) and ε k are the HF spin orbitals and orbital energies, which thus constitute first approximations to the Dyson orbitals and quasiparticle energies. The corresponding noninteracting HF one-particle Green function has the following Lehmann representation[see Eq.(6)] G HF (x 1,x 2 ;ω) = vir a ψ a (x 1 )ψa(x occ 2 ) ω ε a +i0 + + where i and a run over occupied and virtual HF spin orbitals, respectively. i ψ i (x 1 )ψ i (x 2) ω ε i i0 +, (41) Approximations beyond HF are obtained by adding in Eq. (40) a nonlocal energy-dependent correlation self-energy Σ c (x 1,x 2 ;E k ). 8

9 2.2 Linear-response function In the framework of the HF approximation, the IP four-point linear-response function of Eq. (12) becomes the non-interacting HF four-point linear-response function obtained from the non-interacting HF Green function χ HF (1,2;1,2 ) = ig HF (1,2 )G HF (2,1 ), (42) and the IP linear-response function of Eq. (14) becomes the non-interacting HF linear-response function occ virt [ ψ χ HF (x 1,x 2 ;x 1,x 2;ω) = i (x 1 )ψ a(x 1 )ψa(x 2 )ψ i(x 2 ) ω (ε a ε i )+i0 + ψ i (x 2 )ψ a(x 2 )ψa(x 1 )ψ ] i(x 1 ) ω +(ε a ε i ) i0 +. i a Just as the HF self-energy, the HF Bethe-Salpeter kernel has Hartree and exchange contributions Ξ HF Hx(1,2;1,2 ) = Ξ H (1,2;1,2 )+Ξ x (1,2;1,2 ), (44) where the Hartree kernel is found by taking the functional derivatives of Eq. (36) Ξ H (1,2;1,2 ) = i δσ H(1,1 ) δg(2,2) = v(1,2)δ(1,1 )δ(2,2 ) (43) = v(r 1,r 2 )δ(t 1 t 2 )δ(1,1 )δ(2,2 ), (45) and the exchange kernel is found by taking the functional derivatives of Eq. (37) Ξ x (1,2;1,2 ) = i δσ x(1,1 ) δg(2,2) = v(1,2)δ(1,2 )δ(1,2) = v(r 1,r 2 )δ(t 1 t 2 )δ(1,2 )δ(1,2). (46) The presence of the delta functions in Eqs. (45) and (46) implies that the integration over the time variables can easily be performed in the Bethe-Salpeter equation (33), which can then be written for the linear-response function in Fourier space χ TDHF (x 1,x 2 ;x 1,x 2;ω) = χ HF (x 1,x 2 ;x 1,x 2;ω) + dx 3 dx 4 dx 5 dx 6 χ HF (x 1,x 4 ;x 1,x 3 ;ω)ξ HF Hx(x 3,x 6 ;x 4,x 5 )χ TDHF (x 5,x 2 ;x 6,x 2;ω), (47) with the frequency-independent HF Bethe-Salpeter kernel Ξ HF Hx(x 1,x 2 ;x 1,x 2) = v(r 1,r 2 ) [ δ(x 1 x 1)δ(x 2 x 2) δ(x 1 x 2)δ(x 1 x 2 ) ]. (48) Equivalently, Eq. (47) can be written more compactly with inverses (with respect to space-spin coordinates) χ 1 TDHF (x 1,x 2 ;x 1,x 2;ω) = χ 1 HF (x 1,x 2 ;x 1,x 2;ω) Ξ HF Hx(x 1,x 2 ;x 1,x 2). (49) By solving Eq. (47) or (49), we obtain the time-dependent Hartree-Fock (TDHF) linear-response function χ TDHF (x 1,x 2 ;x 1,x 2 ;ω). 9

10 Exercise 8 : Derive Eq. (47) from Eq. (33). 2.3 Linear-response equation in a spin-orbital basis Now, we work out the TDHF linear-response equation in the orthonormal HF spin-orbital basis {ψ k (x)} for practical calculations. One can notice that χ HF (x 1,x 2 ;x 1,x 2 ;ω) in Eq. (43) is expanded into products of two spin orbitals ψp(x 1 )ψ q(x 1 ) and ψ r (x 2 )ψs(x 2 ). The matrix elements of χ HF(x 1,x 2 ;x 1,x 2 ;ω) will thus be labelled by two double indices pq and rs, and are defined by [χ HF (ω)] pq,rs = dx 1 dx 1dx 2 dx 2ψ p (x 1)ψq(x 1 )χ HF (x 1,x 2 ;x 1,x 2;ω)ψr(x 2 )ψ s (x 2). (50) In fact, as seen in Eq. (43), χ HF (x 1,x 2 ;x 1,x 2 ;ω) has non vanishing contributions only on occupied-virtual (ov) and virtual-occupied (vo) products of HF spin orbitals, and is diagonal in this subspace. The matrix representation of its inverse, in the (ov,vo) subspace, is thus immediately found from Eq. (43) to be χ 1 HF (ω) = [( ε 0 0 ε ) ( 1 0 ω 0 1 )], (51) where 1 is the identity matrix and ε is a diagonal matrix of elements ε ia,jb = (ε a ε i )δ ij δ ab, where i,j refer to occupied HF spin-orbitals and a,b to virtual HF spin orbitals. The dimension of the matrix is thus 2M occ M vir 2M occ M vir where M occ and M vir are the numbers of occupied and virtual spin orbitals, respectively. The matrix elements of the HF Bethe-Salpeter kernel of Eq. (48) are defined in the same way as in Eq. (50). We find [Ξ HF Hx] pq,rs = qr ps, (52) where qr ps = qr ps qr sp are the antisymmetrized two-electron integrals, with the definition qr ps = dx 1 dx 2 ψq(x 1 )ψr(x 2 )v(r 1,r 2 )ψ p (x 1 )ψ s (x 2 ). (53) From the Bethe-Salpeter equation (49), we finally find the matrix representation of the inverse of the TDHF linear-response function, in the (ov,vo) subspace, χ 1 TDHF (ω) = [( A B B A with the matrices A and B of elements ) ( 1 0 ω 0 1 )], (54) A ia,jb = ε ia,jb + aj ib, (55a) B ia,jb = ab ij. (55b) The block structure of Eq. (54) is a consequence of the symmetry of the two-electron integrals, qr ps = ps qr. Moreover, the matrix A is Hermitian (i.e., A jb,ia = A ia,jb ) and the matrix B is symmetric (i.e, B jb,ia = B ia,jb ). 10

11 Exercise 9 : Check Eq. (54) and the symmetry properties of the matrices A and B. Similarly to Eq. (11), the TDHF excitation energies ω n are given by the poles of χ TDHF (ω). They correspond to the values of ω giving zero eigenvalues of the matrix χ 1 TDHF (ω) in Eq. (54), leading to the following generalized eigenvalue equation ( )( ) ( )( ) A B Xn 1 0 Xn B A = ω n, (56) 0 1 Y n which is known as the TDHF linear-response equation. The solutions of Eq. (56) come in pairs: if (X n,y n ) is an eigenvector with eigenvalue ω n, then it is easy to check that (Y n,x n) is also an eigenvector with opposite eigenvalue ω n. The positive eigenvalues, ω n > 0, correspond to excitation energies, while the negative eigenvalues, ω n < 0, correspond to de-excitation energies. The 2 2-block matrix on the left-hand-side of Eq. (56) is Hermitian, but the 2 2-block matrix on the right-hand-side of Eq. (56) is not positive definite. Consequently, Eq. (56) is not a Hermitian eigenvalue equation but a pseudo-hermitian eigenvalue equation. One can show that if the 2 2-block matrix on the left-hand-side of Eq. (56) is positive definite then all the eigenvalues ω n are guaranteed to be real numbers (see Exercise 10). Since this matrix corresponds to the Hessian of the HF total energy with respect to the orbital rotation parameters, it is positive definite if the HF solution is a stable minimum with respect to the variation of all orbital parameters. This is known as the Hartree-Fock stability conditions. Y n Exercise 10 : Defining the matrices ( ) A B L = B A and = ( ), (57) Eq. (56) is equivalent to finding the eigenvalues of H = 1 L. 1. Check that L and are Hermitian. 2. Show that H is pseudo-hermitian with respect to L, i.e. H = L H L If L is positive definite, show that H = L 1/2 H L 1/2 is Hermitian. Show then that H has the same eigenvalues than H, and are thus real. For real-valued spin orbitals, and if A B is positive definite, the pseudo-hermitian eigenvalue equation (56) can be transformed into a convenient half-size symmetric eigenvalue equation M Z n = ω 2 n Z n, (58) with M = (A B) 1/2 (A+B) (A B) 1/2 and new eigenvectors Z n. Exercise 11 : Derive Eq. (58) from Eq. (56). Finally, it is common to consider the so-called Tamm-Dancoff approximation (TDA) to linearresponse theory which corresponds to neglecting the matrix B in Eq. (56). After setting B = 0, Eq. (56) simplifies to a standard Hermitian eigenvalue equation A X n = ω n X n. (59) 11

12 With the matrix A in Eq. (55a), Eq. (59) corresponds in fact to the configuration-interaction method with only single excitations (CIS). 2.4 Spin-adapted linear-response equation for closed-shell systems We give now the expressions for spin-restricted closed-shell calculations. We thus consider spatial HF orbitals φ iσ (r) with spin σ = or, so that φ i (r) = φ i (r). The matrices A and B have then each the following spin block structure C = C, C, 0 0 C, C, C, C, 0 0 C, C,, (60) and can be transformed to block-diagonal spin-adapted matrices C = U T CU by the orthogonal transformation U = (61) Applying this transformation to the matrix A gives the following decomposition into singlet and triplet excitation contributions with à = 1 A A A A, (62) 1 A ia,jb = ε ia,jb +2 aj ib aj bi, (63a) 3 A ia,jb = ε ia,jb aj bi. (63b) Similarly, applying the transformation (61) to the matrix B gives with B = 1 B B B B, (64) 1 B ia,jb = 2 ab ij ab ji, (65a) 3 B ia,jb = ab ji. (65b) 12

13 Exercise 12 : Check that the matrices A and B have indeed the form given in Eq. (60), and work out Eqs. (62) and (64). The linear-response eigenvalue equation(56) then decouples into a singlet eigenvalue equation ( 1 A 1 )( B 1 ) ( X n 1 B 1 A 1 = ω Y n n 0 1 and a three-fold degenerate triplet eigenvalue equation ( 3 A 3 )( B 3 ) ( X n 3 B 3 A 3 = ω Y n n 0 1 )( 1 ) X n 1, (66) Y n )( 3 ) X n 3. (67) Y n Note that the minus sign in the last triplet block in Eq. (64) can be absorbed in 3 Y n so that the three triplet eigenvalue equations can be written in a common form. Again, for real orbitals, each eigenvalue equation (66) or (67) can be transformed into a half-size symmetric eigenvalue equation as in Eq. (58). 2.5 Example: H 2 in a minimal basis As a pedagogical example, we consider the calculation of the TDHF excitation energies of the H 2 minimal in a minimal basis consisting of two basis functions, ϕ a (r) and ϕ b (r), centered on each hydrogen atom. This is a closed-shell molecule, therefore all the calculations are done with spin adaptation in a spatial orbital basis. The molecular orbitals are fixed by spatial symmetry φ 1 (r) = ϕ a(r)+ϕ b (r) 2(1+Sab ) (symmetry σ g ), (68) and φ 2 (r) = ϕ a(r) ϕ b (r) 2(1 Sab ) (symmetry σ u ), (69) where S ab is the overlap between ϕ a (r) and ϕ b (r). In HF, the orbital 1 is occupied and the orbital 2 is virtual. The corresponding HF orbital energies are ε 1 = h 11 +J 11, (70) and ε 2 = h 22 +2J 12 K 12, (71) where h 11 and h 22 are one-electron integrals, and J 11 = and K 12 = are twoelectron integrals. The singlet matrices 1 A and 1 B are each one-dimensional 1 A = ε+2k 12 J 12, (72a) 1 B = K 12, (72b) where ε = ε 2 ε 1, leading to the singlet excitation energy 1 ω = ( 1 A+ 1 B)( 1 A 1 B) = ( ε+3k 12 J 12 )( ε+k 12 J 12 ). (73) 13

14 Similarly, the triplet matrices 3 A and 3 B are each one-dimensional 3 A = ε J 12, (74a) leading to the triplet excitation energy 3 B = K 12, (74b) 3 ω = ( 3 A+ 3 B)( 3 A 3 B) = ( ε K 12 J 12 )( ε+k 12 J 12 ). (75) 14

15 Appendices A Fourier-transform formulas We use the following convention for the Fourier transform a function f(τ) f(ω) = and the inverse Fourier transform is thus f(τ) = dτ f(τ) e iωτ, (76) dω 2π f(ω) e iωτ. (77) Here are some useful Fourier-transform formulas involving the Dirac delta function and the Heaviside step function δ(τ) = θ(τ) = θ( τ) = dω 2π e iωτ, (78) dω 2πi dω 2πi e iωτ ω +i0 +, (79) e iωτ ω i0 +. (80) 15

16 B Equation of motion for the one-particle Green function The annihilation field operator in the Heisenberg picture ˆψ(1) = e iĥt 1 ˆψ(x1 )e iĥt 1 obeys the following equation of motion ˆψ(1) = i [ˆψ(1), Ĥ] = ie iĥt 1 [ˆψ(x1 ),Ĥ] e iĥt 1. (81) ] Thecommutator[ˆψ(x1 ),Ĥ is easily calculated from the expression of the Hamiltonian in Eq.(1) using the anticommutation relations of the field operators [ˆψ(x1 ),Ĥ] = h(r 1 )ˆψ(x 1 )+ dx 2ˆψ (x 2 )v(r 1,r 2 )ˆψ(x 2 )ˆψ(x 1 ), (82) where we have used the symmetry of the two-electron interaction, v(r 1,r 2 ) = v(r 2,r 1 ). Introducing now the time exponentials that were in Eq. (81), we find ˆψ(1) = ih(1)ˆψ(1) i d3ˆψ (3)v(1,3)ˆψ(3)ˆψ(1), (83) where we have introduced the notations h(1) = h(r 1 ) and v(1,3) = v(r 1,r 3 )δ(t 1 t 3 ) for convenience. to t 1 We can now calculate the derivative of the one-particle Green function G(1, 2) with respect i G(1,2) = ( ) θ(t 1 t 2 ) N ˆψ(1)ˆψ (2) N θ(t 2 t 1 ) N ˆψ (2)ˆψ(1) N = θ(t 1 t 2 ) N ˆψ(1)ˆψ (2) N +θ(t 1 t 2 ) N ˆψ(1) t ˆψ (2) N 1 θ(t 2 t 1 ) N ˆψ (2)ˆψ(1) N θ(t 2 t 1 ) N ˆψ (2) ˆψ(1) N. (84) Using θ(t 1 t 2 )/ = θ(t 2 t 1 )/ = δ(t 1 t 2 ) and Eq. (83), we find after some simplications i G(1,2) = δ(1,2)+h(1)g(1,2) i d3v(1, 3) [ ] θ(t 1 t 2 ) N ˆψ (3)ˆψ(3)ˆψ(1)ˆψ (2) N θ(t 2 t 1 ) N ˆψ (2)ˆψ (3)ˆψ(3)ˆψ(1) N. (85) The second term can written with the two-particle Green function i 2 G 2 (1,3;2,3 ) = N T[ˆψ(1)ˆψ(3)ˆψ (3 )ˆψ (2)] N. (86) Paying attention to the time ordering and the anticommutations of field operators, we find i G(1,2) = δ(1,2)+h(1)g(1,2) i d3v(1,3)g 2 (1,3 + ;2,3 ++ ), (87) where 3 + means that the time variable has an infinitesimal positive shift t + 3 = t 3+0 +, and 3 ++ means another shift t ++ 3 = t After some rearrangements, we finally write the equation of the motion of G(1,2) as [ i ] h(1) G(1,2)+i d3v(1,3)g 2 (1,3 + ;2,3 ++ ) = δ(1,2). (88) 16

17 C Schwinger functional derivative technique WARNING: this appendix needs to be checked, improved, and extended. See Refs. [1,2,8]. We define a time-dependent Hamiltonian Ĥ(t) = Ĥ +Û(t), (89) where Ĥ is the time-independent many-body Hamiltonian of Eq. (1) and Û(t) is a timedependent external perturbation Û(t 1 ) = dx 1 d1 ˆψ (x 1 )u(1,1 )ˆψ(x 1) (90) whereu(1,1 ) = u(x 1,x 1 ;t 1)δ(t 1 t 1 )isanarbitrarynonlocalpotential, assumedtobehermitian at any t 1, u(x 1,x 1 ;t 1) = u(x 1,x 1;t 1 ), and to vanish sufficiently fast at t 1 ±. The timedependent N-electron state is denoted by N(t) and satisfies the time-dependent Schrödinger equation i N(t) = Ĥ(t) N(t). (91) t C.1 Generalized Green functions and Dyson equation By convention we use the N-electron state at t = 0, N(0), to define the generalized oneparticle Green function ig(1,2) = N(0) T[ˆψ(1)ˆψ (2)] N(0), (92) N(0) N(0) where ˆψ(1) = Û(0,t 1)ˆψ(x 1 )Û(t 1,0) and ˆψ (2) = Û(0,t 2)ˆψ (x 2 )Û(t 2,0) are the annihilation and creation operators in the Heisenberg picture, where Û(t,0) is the time-evolution operator from 0 to t, defined by N(t) = Û(t,0) N(0). From the time-dependent Schrödinger equation (91), it is easy to check that the operator Û(0, t) obeys the evolution equation and, since Û(t,0) = Û(0,t), U(t,0) t Û(0,t) t = iĥ(t)û(t,0), (93) = iû(0,t)ĥ(t). (94) The annhilation operator ˆψ(1) thus satisfies an equation of motion similar to Eq. (81) ˆψ(1) = t iû(0,t 1) [ˆψ(x1 ]Û(t1 ),Ĥ(t),0), (95) 1 which leads to an equation similar to Eq. (83) but with the external potential u(1,1 ) ˆψ(1) = ih(1)ˆψ(1) i d1 u(1,1 )ˆψ(1 ) i d3ˆψ (3)v(1,3)ˆψ(3)ˆψ(1). (96) The same steps used from Eq. (84) to Eq. (88) then give the equation of the motion in the presence of the external potential [ i ] h(1) G(1,2) d1 u(1,1 )G(1,2)+i d3v(1,3)g 2 (1,3 + ;2,3 ++ ) = δ(1,2). (97) 17

18 Repeating the calculations done in Section 1.3, we finally find the Dyson equation satisfied by the generalized one-particle Green function in the presence of the external potential G 1 (1,1 ) = G 1 h (1,1 ) u(1,1 ) Σ Hxc (1,1 ). (98) C.2 Interaction picture and functional derivative We can rewrite the generalized one-particle Green function in the interaction picture ig(1,2) = N T[Ŝ ˆψ I (1)ˆψ I (2)] N N Ŝ N, (99) whereψ I (1) = e iĥt 1 ψ(x 1 )e iĥt 1 andψ I (2) = eiĥt 2 ψ (x 2 )e iĥt 2 aretheannihilationandcreation operators in the interaction picture, and Ŝ is the evolution operator from to in the interaction picture ( ) Ŝ = U I (, ) = T exp i dt 1 Û I (t 1 ), (100) with the interaction-picture perturbation operator ÛI(t 1 ) = e iĥt 1 Û(t 1 )e iĥt 1. In Eq. (99), N is the ground state of the unperturbed Hamiltonian Ĥ which is assumed to be the state at t = ±, so that N(0) = ÛI(0, ) N and N(0) = N ÛI(,0). Using Eq. (90), the operator Ŝ can be written as ( ) Ŝ = T exp i d1d1 ˆψ I (1)u(1,1 )ˆψ I (1 ). (101) Similarly, the generalized two-particle Green function is i 2 G 2 (1,2;1,2 ) = N T[Ŝ ˆψ I (1)ˆψ I (2)ˆψ I (2 )ˆψ I (1 ) N N Ŝ N. (102) The generalized Green functions depend on the potential u only through the operator Ŝ. The functional derivative of G(1,1 ) with respect to u(2,2) is δg(1,1 ) δu(2 = i N T[ δŝ ˆψ δu(2,2) I (1)ˆψ I (1 )] N,2) N Ŝ N N T[Ŝ ˆψ I (1)ˆψ I (1 δŝ )] N N δu(2,2) N. N Ŝ N N Ŝ N The functional derivative of Ŝ is easily obtained from Eq. (101) (103) δŝ δu(2,2) = it[ŝ ˆψ I (2 )ˆψ I (2)] = it[ŝ ˆψ I (2)ˆψ I (2 )], (104) and, when inserted in Eq. (103), leads to δg(1,1 ) δu(2,2) = i2 G 2 (1,2;1,2 ) ig(1,1 )ig(2,2) = iχ(1,2;1,2 ). (105) We thus show that χ(1,2;1,2 ) = i δg(1,1 ) δu(2,2). (106) 18

19 References [1] G. Csanak, H. Taylor and R. Yaris (Academic Press, 1971), vol. 7 of Adv. At. Mol. Phys., p [2] G. Strinati, Rivista del Nuovo Cimento 11, 1 (1988). [3] E. K. U. Gross, E. Runge and O. Heinonen, Many-particle theory (Verlag Adam Hilger, Bristol, 1991). [4] G. Onida, L. Reining and A. Rubio, Rev. Mod. Phys. 74, 601 (2002). [5] F. Bruneval, Ph.D. thesis, Ecole Polytechnique (2005). [6] J. Toulouse, W. Zhu, J. G. Ángyán and A. Savin, Phys. Rev. A 82, (2010). [7] E. Rebolini, J. Toulouse and A. Savin, in Electronic Structure and Reactivity, edited by S. K. Ghosh and P. K. Chattaraj (CRC Press, 2013), Concepts and Methods in Modern Theoretical Chemistry Vol. 1, pp , preprint at [8] F. Aryasetiawan, 19