The GS in presence of vortices

Notice that Ψ⁽⁰⁾₁ ^†e^−iΩ2/2|01i = −Ψ⁽⁰⁾₂ ^†e^−iΩ1/2|10i = e^{−i(Ω1+Ω2)/2}|11i,Ψ⁽⁰⁾₁ |11i = e^−iΩ2/2|01iandΨ⁽⁰⁾₂ |11i=−e^−iΩ1/2|10i.

If the BdG ZM were true fermionic operators, we could generate different orthogo-nal degenerate GS by creating fermions in the localized states with zero energy, as in (6.36). With2nvortices we would then have a GS degeneracy of2²ⁿ, since each ZM could be empty or singly occupied. However this is not the case, and the ZM fields are sums of localized true fermionic creation and annihilation operators. In particular, the states (6.36)are notthe GS, as we will see, but they form a basis on which the GS can be written.

We can construct a different BdG spinor for thek-th vortex, orthogonal by construc-tion to the ZM one, as

B_k⁽⁰⁾= i w_k⁽⁰⁾(r)

−i w⁽⁰⁾_k ^∗(r)

!

, (6.40)

generating a second Majorana operator x⁽⁰⁾_k =

Z drh

i w⁽⁰⁾_k (r)e^iΩk/2ψ(r)−i w⁽⁰⁾_k ^∗(r)e^−iΩk/2ψ^†(r)i

≡ i

√2

Ψ⁽⁰⁾_k −Ψ⁽⁰⁾_k ^†

=x⁽⁰⁾_k ^†. (6.41)

Being orthogonal to A⁽⁰⁾_k , the spinor B_k⁽⁰⁾ can be expanded on all the BdG solu-tions withfiniteenergy (both positive and negative, butnotzero), described by the spinors

S_E,k=

u_E,k(r) v_E,k(r)

. (6.42)

The expansion results in

B_k⁽⁰⁾= i w_k⁽⁰⁾(r)

−i w⁽⁰⁾_k ^∗(r)

!

=X

E6=0

C_E,k⁽¹⁾

u_E,k(r) v_E,k(r)

=X

E6=0

C_E,k⁽¹⁾S_E,k (6.43)

with the coefficients C_E,k⁽¹⁾ =S_E,k^† B⁽⁰⁾_k =

Z drh

i u^∗_E,k(r)w_k⁽⁰⁾(r)−i v_E,k^∗ (r)w⁽⁰⁾_k ^∗(r)i

. (6.44)

Thepositive energypart of B_k⁽⁰⁾ is associated to the annihilationoperator on the 0-generation states. Explicitly, thisfermionicannihilator is generated by

Y_k⁽¹⁾ =X

E>0

C_E,k⁽¹⁾S_E,k= 1 2

B⁽⁰⁾_k −i A⁽¹⁾_k

(6.45) with

A⁽¹⁾_k =iX

E6=0

sgn(E)C_E,k⁽¹⁾S_E,k≡ w⁽¹⁾_k (r) w⁽¹⁾_k ^∗(r)

!

. (6.46)

Notice that, having used the relationv_−E,k=u^∗_E,k, we obtained thatA⁽¹⁾_k is again a Majorana spinor, this time called of ”1-generation”. The 0-generation quasiparticle-excitation annihilator (6.45) can then be written as

y_k⁽¹⁾= i 2√

2

Ψ⁽⁰⁾_k −Ψ⁽⁰⁾_k ^†−Ψ⁽¹⁾_k −Ψ⁽¹⁾_k ^†

(6.47) with obvious notations.

We can already see that the fermionic ZM occupation states on thek-th vortex,|0ik

ande^−iΩk/2|1ik, do not represent two independent GS. In fact, the true GS must be destroyed by (6.47), but if we act with it on the two states above we obtain

y_k⁽¹⁾|0ik =− i 2√

2e^−iΩk/2|1ik y_k⁽¹⁾e^−iΩk/2|1ik = i 2√

2|0ik . (6.48) Thus, we can already predict that the true GS will imply entanglement of the fermionic ZM occupation states.

At this point we can iterate the arguments. Out ofA⁽¹⁾_k we can construct the orthog-onal Majorana spinor

B_k⁽¹⁾= i w⁽¹⁾_k (r)

−i w⁽¹⁾_k ^∗(r)

!

= X

E6=0

C_E,k⁽²⁾S_E,k (6.49) with

C_E,k⁽²⁾ =S_E,k^† B⁽¹⁾_k . (6.50) Thepositive energypart ofB_k⁽¹⁾generates theannihilationoperator on the 1-generation states

Y_k⁽²⁾ =X

E>0

C_E,k⁽²⁾S_E,k= 1 2

B⁽¹⁾_k −i A⁽²⁾_k

(6.51) with

A⁽²⁾_k =iX

E6=0

sgn(E)C_E,k⁽²⁾S_E,k≡ w_k⁽²⁾(r) w_k^(2)∗(r)

!

(6.52) and so on for successive generations.

Thej-generation localized functionsw^(j)(r)are generic, until the BdG functions in S_E,k are obtained explicitly. Still, it is possible to show generally (see Appendix B) thateachw^(j)(r)is orthogonal toevery otherw^(j0)(r)ifj 6=j⁰. This implies that all theA^(j)_k andB^(j)_k spinors are orthogonal to one another if they belong to different generations, apart from being trivially orthogonal, due to their localization, if they sit on different vortices.

This iteration automatically produces orthogonal localized wavefunctions that can be used as a functional basis around each vortex (completeness is still to be proved, but we believe it holds in the case of infinitely far vortices). The successive gen-eration essentially stops whenever thewfunctions start overlapping with those of other vortices.

Let me express my deep appreciation to Ady Stern for inventing this beautiful pro-cedure.

Apart from its clean logical beauty, the generation procedure presented above turns out to be extremely suitable to address the nature of the GS in the many vortices case.

In order to produce azero energytrue fermionic field withtwovortices we con-sider the operator

α⁽⁰⁾=α⁽⁰⁾₁ −i α⁽⁰⁾₂ = 1

√2

Ψ⁽⁰⁾₁ + Ψ⁽⁰⁾₁ ^†−iΨ⁽⁰⁾₂ −iΨ⁽⁰⁾₂ ^†

. (6.53) In complete analogy we can build

x⁽⁰⁾=x⁽⁰⁾₁ −i x⁽⁰⁾₂ = 1

√2

iΨ⁽⁰⁾₁ −iΨ⁽⁰⁾₁ ^†+ Ψ⁽⁰⁾₂ −Ψ⁽⁰⁾₂ ^†

(6.54) and introduce four orthogonal states

| ↓↓i (6.55)

| ↑↓i=α^(0)†| ↓↓i (6.56)

| ↓↑i=x^(0)†| ↓↓i (6.57)

| ↑↑i ≡α^(0)†x^(0)†| ↓↓i (6.58) with the conditions

α⁽⁰⁾| ↓↓i=x⁽⁰⁾| ↓↓i= 0. (6.59) To get the relation between this ”spin” description and the occupation representa-tion we write the state| ↓↓ion the basis (6.36) as

| ↓↓i=a|00i+b e^−iΩ2/2|01i+c e^−iΩ1/2|10i+d e^{−i(Ω1+Ω2)/2}|11i. (6.60) Imposing (6.59) with the definitions (6.53,6.54) we get the conditionsa=d= 0and c=i·b, producing

| ↓↓i= 1

√2

e^−iΩ1/2|10i −i e^−iΩ2/2|01i

(6.61) where normalization has been implemented.

The successive terms in (6.55) give

| ↓↑i= 1

√2

|00i+i e^{−i(Ω1+Ω2)/2}|11i

| ↑↓i= 1

√2

|00i −i e^{−i(Ω1+Ω2)/2}|11i

(6.62)

| ↑↑i= 1

√2

e^−iΩ1/2|10i+i e^−iΩ2/2|01i .

Altogether, we see that each spin-description state implies entanglement in the oc-cupation representation. Notice that no tunneling is ever considered. The entan-glement holds for states on different vortices and is therefore non-local, somehow

resembling what happens in the EPR (Einstein-Podolski-Rosen) paradox.

As discussed qualitatively above, for two vortices we expecttwo GS. In the spin-representation they are easily identified as the two possible states for the zero-energy operatorα⁽⁰⁾. By tracing over the ”right” spin states and stopping at the 0-generation level, we obtain the two possibilities

| ↓i=| ↓↓i+| ↓↑i (6.63)

| ↑i=| ↑↓i+| ↑↑i. (6.64) Using the (6.61,6.62) we can finally express the two GS in the occupation represen-tation. The entanglement between states of the same generation is evident, and we observe that each GS is a superposition of many-body configurations with different particle numbers, as we already saw for the BCS state (chapter 4).

Before characterizing completely the structure of the GS for the next generations, we can already address the issue of non-Abelian quasiparticle statistics. The ques-tion is, as seen in chapter 1, what happens when one of the two vortices encircles the other.

We can for instance drag vortex 2 adiabatically in a closed loop around vortex 1 (which can be chosen to sit in the origin). The occupation of the ZM is not affected by this operation, but the phase Ω₂ goes intoΩ₂ + 2π. The final effect is thus a change of sign for the states with vortex 2 occupied in the family (6.61,6.62). There-fore, in (6.63) the GS| ↓iis transformed into| ↑iand viceversa. These statistical factors coincide with those obtained in [135, 140].

Our derivation offers a quite physical picture of the non-Abelian statistics. The in-gredients which have been shown to be crucial in our discussion are the presence of ZM in the vortex cores leading to the GS degeneracy, the entanglement between states living on different vortices and the phase accumulated by a vortex dragged around a closed loop in the 2DES.

Having obtained the mapping between the occupation and ”spin” representa-tions, we can now procede to determine the full structure of the GS. In particular, once we solved the issue of entanglement at the 0-generation level, we want to ad-dress what happens to the further generations.

We saw that, while building the quasiparticle-excitation annihilators on the j-th generationy_k^(j+1), we automatically produced the Majorana spinor ofj+ 1 genera-tionA^(j+1)_k . The annihilators will therefore constitute the bridges between different generations.

To see this, we need a two-spins ketper each generationj, of the form|s_j, S_jiwith s_j, S_j =↑/↓, such thatα^(j)(x^(j)) lowers the spin values_j(S_j), exactly as in (6.55).

The two GS in the two-vortices case are therefore identified withs⁽⁰⁾ =↑,↓, as in (6.63). To keep trace of the next generations we can (tensor)-multiply each two-spin ket|s_j, S_jiwith a ”bath” of states describing the further generations, indicated as

BS^(j+1)_j . In this notation we write the two GS as

| ↓i=| ↓↓i|B⁽¹⁾_↓ i+| ↓↑i|B_↑⁽¹⁾i (6.65)

| ↑i=| ↑↓i|B⁽¹⁾_↓ i+| ↑↑i|B_↑⁽¹⁾i (6.66) or, more compactly

|s₀i=X

S₀

|s0, S₀i|BS⁽¹⁾₀i. (6.67) At this point we impose that the GS are destroyed by the annihilatorsy⁽¹⁾_k (6.47).

More conveniently we can define the operators

y^(j)_± =y₁^(j)±iy₂^(j) (6.68) with

y^(j)₊ =1 2

x^(j−1)†−iα^(j)†

(6.69) y^(j)₋ =1

2

x^(j−1)−iα^(j)

. (6.70)

By requesting

y₊⁽¹⁾|s₀i= 0 (6.71)

y₋⁽¹⁾|s₀i= 0 (6.72)

fors₀ =↑,↓, and bracketing (6.71) withhs₀, S₀|for everyS₀ =↑,↓, we obtain four equations for the baths

α^(1)†|B⁽¹⁾S₀=↑i=i|BS⁽¹⁾₀=↓i (6.73) α^(1)†|B⁽¹⁾S₀=↓i= 0 (6.74)

α⁽¹⁾|BS⁽¹⁾₀=↑i= 0 (6.75)

α⁽¹⁾|BS⁽¹⁾₀=↓i=−i|B⁽¹⁾S₀=↑i. (6.76) Therefore the bath|BS⁽¹⁾₀ihas the typical behaviour of the spin state withs₁ =−S₀ and can then be written as

|B⁽¹⁾S₀i=i^S0/2X

S₁

|s₁=−S₀, S₁i|B⁽²⁾S₁i (6.77) where we associateS_j=↓≡ −1andS_j=↑≡1(same fors_j). Analogous arguments can be repeated for baths of successive generations.

The GS have therefore a beautiful self-similar structure. Along with the vortex-entanglement at the same generation level implied in the two-spins kets|s_j, S_ji(see

(6.55)) the (6.71) also produceentanglement between neighbour generations. The final form of the GS will therefore be

|s₀i = X

S₀

|s₀, S₀ii^S0/2X

S₁

|s₁=−S₀, S₁ii^S1/2X

S₂

|s₂=−S₁, S₂ii^S2/2×

×X

S₃

... i^Sn/2 X

S_n+1

|s_n+1=−S_n, S_n+1i|B⁽ⁿ⁺²⁾S_n+1 i. (6.78) The indexnindicates the upper limit of the generation procedure, occurring when thew_k⁽ⁿ⁾functions on different vortices start overlapping.

In the case of four (and more) vortices the arguments are quite analogous. The vor-tices will be divided in many pairs. Within each pair the two ZM will be combined to form a single complex true-fermionic operator, and the successive generation procedure works as already presented.

The analysis presented up to now is completely general and highlights the complex structure of the GS.

We could however dig deeper into the knowledge of the properties of our system if we succeeded in solving the spectrum of spinor BdG eigenstates. This would allow to address the localized statesw_k^(j)explicitly and would produce a different basis on which to expand the paired wavefunctions.

Indeed, one aspect which is still to be solved is the issue of pairing in the inhomo-geneous p-wave state. That is, we would now like to determine what is the nature of the Cooper pairs wavefunctions, i.e. address the open question ”who pairs to whom?”.

In order to do that we will now consider a case in which the BdG equations can be solved exactly and finally write the GS on the basis of the BdG states.

Im Dokument On the role of spin, pairing and statistics for composite fermions in the fractional quantum hall effect (Seite 154-160)