Abstract
Connecting double quantum dots via a semiconductorsuperconductor hybrid segment offers a platform for creating a twosite Kitaev chain that hosts Majorana zero modes at a finely tuned sweet spot. However, the effective couplings mediated by Andreev bound states in the hybrid are generally weak in the tunneling regime. As a consequence, the excitation gap is limited in size, presenting a formidable challenge for using this platform to demonstrate nonAbelian statistics and realize topological quantum computing. Here we systematically study the effects of increasing the dothybrid coupling. In particular, the proximity effect transforms the dot orbitals into YuShibaRusinov states, and as the coupling strength increases, the excitation gap is significantly enhanced and sensitivity to local perturbation is reduced. We also discuss how the strongcoupling regime shows in experimentally accessible quantities, such as conductance, and provide a protocol for tuning a doubledot system into a sweet spot with a large excitation gap.
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Introduction
The Kitaev chain is a toy model of topological superconductivity that consists of onedimensional spinless fermions with pwave pairing potential^{1}. In the topological phase, the endpoints host a pair of Majorana zero modes^{2,3,4,5,6,7,8,9,10,11,12,13}, which obey nonAbelian statistics and are regarded as the building block of topological quantum computation^{14,15}. Such a Majorana qubit is predicted to be more immune to decoherence due to the quantum information being encoded nonlocally in space and further protected by an excitation gap above the computational subspace.
In solidstate physics, the Kitaev chain model can be simulated in a quantum dot array by utilizing the spinpolarized dot orbitals as spinless fermions, with the effective couplings mediated by superconductivity^{16}. Remarkably, even a chain consisting of only two quantum dots can exhibit finetuned, but still spatially separated Majorana modes at a sweet spot, colloquially called poor man’s Majorana modes^{17}. Recently, such a twosite Kitaev chain was experimentally realized in double quantum dots, and poor man’s Majorana modes were identified via conductance spectroscopy at the sweet spot^{18}. In particular, the effective couplings, both normal and superconducting ones, are mediated by an Andreev bound state (ABS) in a hybrid segment connecting both quantum dots^{19}, which allows for a deterministic finetuning of the relative amplitude by changing the ABS chemical potential via electrostatic gating^{20,21,22}. This effect was shown theoretically to be robust to Coulomb interactions in the dots as well as stronger coupling^{23}.
Despite the experimental progress, stateoftheart Kitaev chain devices are still constrained by a relatively small excitation gap (~25 μeV), which is much smaller than the induced gap of the ABS (~150 μeV) and the parent aluminum gap (~230 μeV)^{18}. In order to experimentally demonstrate the nonAbelian statistics of Majoranas and to obtain highquality Majorana qubits^{24,25,26}, a significant enhancement in the excitation gap is crucial. This enhancement will allow for a more tolerant adiabatic limit condition \(\sim \hslash /{E}_{{{{{{{{\rm{gap}}}}}}}}}^{1}\)^{27,28,29,30} and suppress the detrimental thermal effects \(\sim {e}^{{E}_{{{{{{{{\rm{gap}}}}}}}}}/{k}_{B}T}\)^{31}.
In this work, we use the threesite model^{19,23,32} to systematically study enhancing the energy gap by increasing the dothybrid coupling strength, achievable in experiments by lowering the tunnel barrier height. As a result of the proximity effect from the hybrid, the spinpolarized orbitals in the quantum dots undergo a transformation into Yu–Shiba–Rusinov (YSR) states^{33,34}, in an analogy with the conventional YSR states^{35,36,37}. These states then constitute the new spinless fermion basis for the emulated Kitaev chain. Thus, the concepts of elastic cotunneling and crossed Andreev reflection in the weak coupling regime have to be generalized. Most importantly, we show that poor man’s Majorana zero modes can survive in this strong coupling regime, featuring a significantly enhanced excitation gap. The properties of the resulting states are different from those in the weak coupling regime, showing both wavefunction profiles and conductance properties while maintaining their Majorana character.
Methods
Model and Hamiltonian
A twosite Kitaev chain device consists of two separated quantum dots connected by a hybrid segment [see Fig. 1a]. The system Hamiltonian is^{19,23,32}
where H_{D} is the Hamiltonian of the quantum dots, \({n}_{Da\sigma }={c}_{Da\sigma }^{{{{\dagger}}} }{c}_{Da\sigma }\) is the electron occupancy number on dot a, ε_{Da} is the orbital energy, E_{ZDa} is the strength of the induced Zeeman energy, and U_{Da} is the Coulomb repulsion strength. H_{S} describes the hybrid segment hosting a pair of ABSs in the lowenergy approximation. ε_{A} is the normalstate energy, and Δ_{0} is the induced pairing gap. While we assume no induced Zeeman energy in the ABS due to a strong renormalization effect at the hybrid interface^{38,39}, the main conclusions remain valid for finite Zeeman energy as well. H_{T} is the tunnel coupling between dot and ABS, including both spinconserving ~ t and spinflipping ~ t_{so} processes. In realistic devices, the amplitude of t is a variable that can be controlled by tunnel barrier gates, while the ratio of t_{so}/t is generally fixed and is determined by the strength of spinorbit interaction. In the rest of this work, we will choose Δ_{0} to be the natural unit. Unless stated otherwise, we set E_{ZDa} = 1.5 Δ_{0}, U_{Da} = 5 Δ_{0}, and t_{aso}/t_{a} = 0.3 according to the recent experimental measurements on similar devices^{18,20,21,22}. In addition, we numerically calculate the differential conductance using the rateequation method^{23}, where the lead tunneling rate is Γ_{a} = 0.025Δ_{0}, and temperature is k_{B}T = 0.02 Δ_{0}.
Results
Quantum dotAndreev bound state pair
To assess the strength and to understand the effects of dothybrid coupling, we first focus on the conductance spectroscopy of a single quantum dotABS pair. Hence, for the discussions here, we temporarily remove the right dot in the model Hamiltonian in Eq. (1). Figure 2 shows the zerobias conductance spectroscopy in the (ε_{D}, ε_{A}) plane for t/Δ_{0} = 0.25, 1 and 2, respectively. Here G_{LL} = dI_{L}/dV_{L} at V_{L} = 0. As shown in Fig. 2a, in the weak coupling regime, the conductance resonances are two straight lines extending along ε_{A}, corresponding to the spinup and down orbitals in the quantum dot. In contrast, with a strong dothybrid coupling, the resonance lines become Sshaped curves [see Fig. 2b], and the conductance magnitudes are increased by order of magnitude. For even stronger coupling, Fig. 2c, proximity from the ABS is so strong that the states are very different from spinup and spindown, as signified by the differently connected arcs in the conductance.
This qualitative behavior was already described in existing literature^{34}. Here, we recover the same behavior in a simpler model and also consider the magnitude of the conductance. In particular, we can use secondorder perturbation theory to qualitatively understand the physical mechanisms underlying these conductance features.
First, the dothybrid coupling renormalizes the dot orbital energy by δε_{D} via cotunneling processes. Up to the leading order of t and t_{so}, this energy renormalization is
where u^{2} = 1 − v^{2} = 1/2 + ε_{A}/2E_{A} are the BCS coherence factors and \({E}_{A}=\sqrt{{\varepsilon }_{A}^{2}+{\Delta }_{0}^{2}}\) is the ABS excitation energy. Interestingly, due to destructive interference, the dot energy shifts positively (negatively) for ε_{A} > 0 (ε_{A} < 0), vanishes at ε_{A} = 0, and decays as \({\varepsilon }_{A}^{1}\) for large ε_{A}, well explaining the Sshaped conductance resonances shown in Fig. 2b. Hence, the Sshaped feature is a clear sign of the proximity effect due to the ABS in the hybrid semiconductor–superconductor segment (rather than directly to the parent metallic superconductor), and the increase in bending is a direct signal of increasing coupling.
From Fig. 2a, b, we also observe that conductance is largest when the ABS is near resonance, i.e., −Δ_{0} < ε_{A} < Δ_{0}, which can also be understood from perturbation theory. The ABS will induce a pairing term on the quantum dot, i.e., \({\Delta }_{{{{{{{{\rm{ind}}}}}}}}}{c}_{D\uparrow }^{{{{\dagger}}} }{c}_{D\downarrow }^{{{{\dagger}}} }+h.c.\) via local Andreev reflection with
The induced gap size is prominent within −Δ_{0} < ε_{A} < Δ_{0} and decays as \({\varepsilon }_{A}^{2}\) outside. As a result, the local Andreev conductance is significantly enhanced when ABS is near resonance and vanishes when offresonance. Thus, based on the shape of the resonance lines as well as on the enhancement of the Andreev conductance, one can estimate the strength of the dothybrid coupling in an actual device. In addition, although not the main focus of the current work, it is likely that the continuum states of the parent superconductor also induce local proximity effect on the quantum dot. Such contribution would be a constant that is independent of the ABS chemical potential, i.e., \({\Delta }_{{{{{{{{\rm{ind}}}}}}}}}\to {\Delta }_{{{{{{{{\rm{ind}}}}}}}}}+{\Delta }_{{{{{{{{\rm{ind}}}}}}}}}^{{{{{{{{\rm{qp}}}}}}}}}\). More importantly, especially in the strong coupling regime, the induced superconducting effect would transform the dot orbitals into YSR states^{34,40,41,42,43,44,45,46}. These states then establish the new spinless fermion basis for the emulated Kitaev chain.
Coupled YSR states
We now turn to the case of two quantum dots coupled via an ABS and develop an effective theory for two coupled YSR states. By assuming that the ABS in the hybrid remains gapped, we can integrate it and obtain the effective coupling
where t_{ση} and Δ_{ση} are the elastic cotunneling (ECT) and crossed Andreev reflection (CAR) amplitudes between electron or hole excitations in the two dots. These couplings are tunable by changing the energy of the middle dot, ε_{A}^{19}. Note that the problem of coupling two quantum dots via ECT and CAR, and in the presence of local Andreev reflection giving rise to a proximity effect in the dots has been studied extensively before^{47,48,49,50}, predominantly at zero magnetic field. In contrast, we focus on the case of a significant Zeeman splitting in the outer quantum dots, such that the ground state of both dots occupied by a single electron is a triplet state. We also emphasize that the treatment in Eq. (4) is valid when the dotABS coupling is weak (t ≪ Δ_{0}) or intermediate in strength (t ≲ Δ_{0}), while the quantum dot orbitals can be strongly proximitized by the ABS or the continuum states in the parent superconductor.
For unproximitized quantum dots, Eq. (4) plus H_{D} in Eq. (1) indeed resembles the Hamiltonian of a twosite Kitaev chain^{1}. However, since YSR states are a superposition of electron and hole components, the effective couplings of ECT and CAR have to be generalized. In particular, for a single proximitized quantum dot with finite Zeeman splitting, the ground states in the even and oddparity subspaces are a spin singlet and a spindown state, respectively,
where u^{2} = 1 − v^{2} = 1/2 + ξ/2E_{0} are the BCS coherence factors, with ξ = ε + U/2 and \({E}_{0}=\sqrt{{\xi }^{2}+{\Delta }_{{{{{{{{\rm{ind}}}}}}}}}^{2}}\), and \(\vert {n}_{\uparrow }{n}_{\downarrow }\rangle\) is a state in the occupancy representation. Consequently, we define YSR state as \(\left\vert \downarrow \right\rangle ={f}_{{{{{{{{\rm{YSR}}}}}}}}}^{{{{\dagger}}} }\left\vert S\right\rangle\), with an excitation energy δε = E_{↓} − E_{S}. When coupling two YSR states via the ABS, the effective Hamiltonian becomes
where f_{a} denotes the YSR state in dota, with δε_{a} being the excitation energy. This also takes the form of a Kitaev Hamiltonian, but now in the basis of YSR states. Crucially, Γ_{o/e} represents the generalized effective couplings between YSR states. The oddparity coupling is
where \(\vert \downarrow S\rangle ,\vert S\downarrow \rangle\) are the tensor states with total parity odd, and u_{a}, v_{a} are the BCS factors defined in Eq. (5). Note that Γ_{o} is a linear combination of equalspin ECT and oppositespin CAR, which are all spinconserving processes. On the other hand, the evenparity coupling is
which couples states with total spin zero and one. In particular, a finite Γ_{e} requires a physical mechanism to break spin conservation, e.g., spinorbit interaction.
Figure 3a shows a schematic of the charge stability diagram as a function of the quantum dot energies. Blue and red squares indicate whether the ground state of two uncoupled dots is odd or even, respectively. Note that the singlet ground state in the dots does depend on the dot energies through the values of u_{a} and v_{a}. For example, in the upperright corner u_{L/R} > v_{L/R}, corresponding to each dot predominantly empty, whereas the lowerleft corner features u_{L/R} < v_{L/R}, corresponding to each dot predominantly doubly occupied. The arrows represent the interactions Γ_{o/e}, and the relative strength of these couplings will determine the ground state close to the four corners in the charge stability diagram where each dot exhibits a degeneracy without interactions. Additionally, we point out the role of CAR and ECT interchange for different corners of the charge stability diagram.
In the absence of spin–orbit interaction, Γ_{e} = 0, and in general Γ_{o} ≠ 0. Hence, at the four corners, the energy of the odd ground state is lowered compared to the even ones. This can also be observed in a simulation of the full threedot Hamiltonian in Eq. (1), as shown in Fig. 3b, where we find a disconnected even island in the center of the charge stability diagram. Note that such behavior is only possible for a finite Zeeman splitting and, as such, qualitatively different from the charge stability diagrams^{49}.
For a system with finite spin–orbit coupling, in general, also Γ_{e} ≠ 0, and the respective values will depend on details of the system (such as the energy ε_{A} of the middle dot that can be used to tune ECT and CAR). In particular, it is now, in general, possible to change the relative strength of Γ_{e/o}. This shows as a change in connectivity in the charge stability diagram, with a guaranteed sweet spot Γ_{e} = Γ_{o} in between. We show this behavior in Fig. 3c–e on the example of the upperright corner as we vary ε_{a}, transitioning from a regime dominated by Γ_{o} in Fig. 3c to one dominated by Γ_{e} in Fig. 3e. When Γ_{e} = Γ_{o}, a cross emerges in the phase diagram as a signature of the sweet spot, as shown in Fig. 3d.
In the limit of large Coulomb interaction U on the outer dots, either u_{L/R} ≈ 1 or v_{L/R} ≈ 1 and Γ_{e/o} will be dominated by a single ECT or CAR term, as evident from Eqs. (7) and (8).
More generally, however, in the limit of vanishing Zeeman splitting in the middle dot, CAR and ECT coupling are constrained by t_{↑↑} = t_{↓↓}, t_{↑↓} = − t_{↓↑}, Δ_{↑↑} = Δ_{↓↓}, Δ_{↑↓} = − Δ_{↓↑}, due to timereversal symmetry, such that Γ_{o/e} can be further simplified as
where 0 ≤ β_{a} ≤ π/2 is a parameter to characterize the BCS factors by \({u}_{a}=\cos {\beta }_{a},{v}_{a}=\sin {\beta }_{a}\). For two dots with a similar level of proximity effect, the diagonal corners in Fig. 3 will have β_{L} = β + δ/2, β_{R} = β − δ/2, with δ ≪ 1 characterizing a weak asymmetry. As a result, the odd and evenparity couplings reduce to
This indicates that as the proximity effect increases, the initially purely equalspin ECT/CAR coupling ratio Γ_{o}/Γ_{e} gains a finite oppositespin CAR/ECT component. In contrast, around the offdiagonal corners, we have β_{L} = β + δ/2, π/2 − β_{R} = β − δ/2, yielding
Interestingly, despite the proximity effect, Γ_{o/e} is equal to the only ECT or CAR just as in the unproximitized regime. Only an asymmetry in the proximity effect leads to a mixing of CAR and ECTtype couplings. For a detailed investigation of how Γ_{e} and Γ_{o} behave in different corners of the charge stability diagram, we refer to Supplementary Note 1: Γ_{e} and Γ_{o} for different corners of the charge stability diagram.
Poor man’s Majorana
We now focus on the properties of the poor man’s Majoranas that appear at the sweet spot in the full dothybriddot system. Without loss of generality, we assume that the left and right dots have the same set of physical parameters, e.g., ε_{DL} = ε_{DR} = ε_{D}, E_{ZL} = E_{ZR} = 1.5Δ_{0}, U_{L} = U_{R} = 5Δ_{0}, t_{L} = t_{R} = t and t_{Lso} = t_{Rso} = 0.3t. To simplify, we introduce a shift in dot energy ε_{D} → ε_{D} − E_{ZD} to set the zero energy of the spindown orbital at ε_{D} = 0. Figure 4a shows the phase diagram in the (ε_{A}, ε_{D}) plane for weakly coupled quantum dots (t/Δ_{0} = 0.25), with \(\delta E={E}_{{{{{{{{\rm{gs}}}}}}}}}^{{{{{{{{\rm{odd}}}}}}}}}{E}_{{{{{{{{\rm{gs}}}}}}}}}^{{{{{{{{\rm{even}}}}}}}}}\) being the energy difference of the ground states in the opposite fermion parity subspace. The whitecolored curves (δE = 0) represent the ground state degeneracy, with the tip of the curves (marked by a black cross sign) indicating the sweet spot^{23}. At this sweet spot, the effective normal and superconducting couplings of the two dots become equal in strength. The wavefunction profiles plotted in Fig. 4b further demonstrate that the Majorana zero modes are well localized at the two dots, respectively, with only a negligible amount of overlap in the middle ABS. Here, the Majorana wavefunction densities are defined for site and spin
where a = L, M, R, σ = ↑, ↓, and \(\left\vert e\right\rangle ,\left\vert o\right\rangle\) denote the even and oddparity ground state.
Comparatively, the lower panels in Fig. 4 show the results obtained in the strong dotABS coupling regime with t/Δ_{0} = 1. Around the sweet spot marked by the black cross sign, the ground state degeneracy line now becomes much broader and straighter compared to the weak coupling regime, indicating a significantly enhanced energy gap and robustness against dot chemical potential fluctuations. In Fig. 4d, the plotted Majorana wavefunctions show strong leakage into the middle ABS and small leakage to the opposite normal dot with opposite spin. We emphasize that the wavefunction overlap in the ABS, which is a virtual state, is not detrimental, and that the reduced density on the normal dots will only reduce the visibility of the Majorana from the external detecting system.
To gain a better understanding of the effect of strong coupling, we now investigate the continuous evolution of the sweet spot and the properties of Majoranas as a function of t. To that end, we define the following quantities: excitation gap (E_{gap}), Majorana localization (ρ), and polarization (χ)^{23} as below
Here, E_{gap} represents the excitation gap above the poor man’s Majorana zero modes, where \(\delta {E}_{oe}^{1}\) and \(\delta {E}_{eo}^{1}\) denote the energy differences between the ground state in one parity sector and the first excited state in the opposite one. The Majorana localization ρ and Majorana polarization χ, both of which are defined on the outer quantum dot, characterize the localization and overlap of the Majorana wavefunction on the normal dot, respectively.
Figure 5a shows the evolution of the positions of the sweet spots in the (ε_{A}, ε_{D}) plane. As the coupling strength t increases, the effect of dot energy renormalization predicted by Eq. (2) becomes more pronounced, making ε_{D} deviate from the value of ε_{D} = 0 in the weak coupling regime. At the same time, the sweetspot values of ε_{A} shift towards a more positive value, indicating an induced Zeeman energy in ABS^{19}, which comes from the inverse proximity effect from the quantum dot. One crucial aspect of the strong coupling regime is that, with increasing t, the excitation gap is enhanced significantly in a nearly quadratic manner [see Fig. 5b]. The excitation gap reaches as high as E_{gap} ~ 0.7Δ_{0} for a dothybrid coupling t/Δ_{0} = 1.1. Moreover, we observe that the degree of protection against detuning of both quantum dots away from the sweet spot increases with the growing t, as evidenced by the diminishing curvature of the quadratic splitting of ground state degeneracy in Fig. 5c. On the other hand, the Majorana localization ρ is largely reduced due to increased wavefunction leakage into the middle ABS when the tunnel barrier is lowered. Yet, although the middle dot hybridizes strongly with the Majoranas in the outer dots, the Majorana polarization χ decreases much slower. Hence, even for strong coupling, the overlap of Majoranas on the outer dots remains small, which will be beneficial for future qubit experiments.
We observe these changes manifesting in the local conductance profile calculated at the sweet spot [see Fig. 5d]. Specifically, as t increases, the height of the Majoranainduced zerobias conductance peaks decreases due to the reduction in Majorana density at the outer dot, which in turn reduces its effective coupling strength with the normal lead. In addition, the voltage bias values where the side peaks appear, indicating the magnitude of the excitation gap, increases with t, and a single side peak begins to split into two at larger t values, consistent with our findings in Fig. 5b.
Signatures of strong coupling in nonlocal transport
Nonlocal transport is a useful tool for probing hybrid systems, as the nonlocal conductances G_{LR} and G_{RL} can measure the BCS charge of states^{51} in noninteracting systems. For example, in a recent experiment, nonlocal conductance was used to confirm the chargeless Majorana character of the zeroenergy state^{18}.
We now discuss how the visibility of the first excited state in the nonlocal conductance G_{LR} is a qualitative indicator of the strong coupling regime. In Fig. 6, we show the nonlocal conductances in both weak coupling (t = 0.25 Δ_{0}) and strong coupling (t = Δ_{0}) regimes, as a function of the applied voltage bias V_{bias} and chemical potential detuning away from the sweet spot δε_{DR}. In the weak coupling regime, shown in Fig. 6a, b, the conductance signal strength for G_{LR} is significantly lower than its counterpart G_{RL}. This behavior is in line with the results for a spinless Kitaev chain model^{18}. There, the chargeless nature of the excited state leads to a vanishing G_{LR} signal. In contrast, as we increase the dothybrid coupling, this effective picture breaks down, and the nonlocal conductances G_{LR} and G_{RL} become comparable in strength, as demonstrated in Fig. 6c, d. Furthermore, G_{RL} for the first excited state changes sign as a function of detuning δε_{DR} in the right dot, whereas G_{LR} does not. The calculations of a complete conductance matrix, including both local and nonlocal conductances, can be found in Supplementary Note 2: Conductance matrix for weak and strong coupling regimes.
In Fig. 6e, we track the evolution of maximum nonlocal conductance signal strength with varying t. We observe that the maximum conductance for G_{LR} increases much faster with increasing t compared to G_{RL}. We can attribute the increase in G_{LR} to the increase in the BCS charge of the excited state in the left quantum dot with increasing t, as shown in Fig. 6f^{51}.
Discussion
In this work, we have studied a dothybriddot system in the strong dothybrid coupling regime (t ~ Δ_{0}) for implementing a twosite Kitaev chain. Due to the proximity effect from the ABS, the dot orbitals undergo a transformation into YSR states, which constitute the new spinless fermion basis for the effective Kitaev chain, and we have studied their coupling as mediated by the middle dot. Importantly, poor man’s Majorana zero modes persist in this strong coupling regime and now possess a significantly enhanced excitation gap. On the other hand, there is an upper bound for the dothybrid coupling strength. As shown in Fig. 2c, an excessively strong coupling t ~ 2E_{ZD} leads to the hybridization of dot orbitals with opposing spins, gapping out the zeroenergy YSR states. That is, a sufficiently strong Zeeman field (E_{ZD} ≫ t) is always a prerequisite for obtaining an effectively spinless Kitaev chain model in a spinful physical system^{52,53}.
Additionally, our theoretical work provides a practical recipe for implementing a twosite Kitaev chain with an enhanced excitation gap. Specifically, one can reach the desirable dothybrid coupling regime by observing the conductance spectroscopy of a single quantum dotABS while shifting the other dot offresonance, e.g., Fig. 2b. Performing this procedure for both left and right pairs, the sweet spot can be further finetuned by changing the chemical potential of the ABS to find a crossing in the charge stability diagram [see Fig. 3c–e] and a robust zerobias peak in the conductance spectroscopy [see Fig. 5d]. Indeed, a parallel experimental work^{54} has achieved an energy gap of approximately ~75 μeV using the aforementioned procedure, verifying the theoretical predictions proposed in this work. We thus expect that our work provides useful guidelines and insights for realizing Kitaev chains in the strong coupling regime, serving as the central platform for future research on Majorana quasiparticles and nonAbelian statistics.
Data availability
The data that support the findings of this study have been deposited in Zenodo^{55}.
Code availability
The code used to generate the figures is available on Zenodo^{55}.
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Acknowledgements
This work was supported by a subsidy for top consortia for knowledge and innovation (TKl toeslag), by the Dutch Organization for Scientific Research (NWO), by the European Union’s Horizon 2020 research and innovation program FETOpen Grant No. 828948 (AndQC), and by Microsoft Corporation. AMB acknowledges NWO (HOTNANO) for the research funding.
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C.X.L. designed the project with input from F.Z., S.L.D.t.H., T.D., and M.W. C.X.L., and A.M.B. performed the calculations and generated the figures. All authors contributed to the discussions and interpretations of the results. C.X.L. and M.W. supervised the project. C.X.L., A.M.B., and M.W. wrote the paper with input from F.Z., S.L.D.t.H., and T.D.
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Liu, CX., Bozkurt, A.M., Zatelli, F. et al. Enhancing the excitation gap of a quantumdotbased Kitaev chain. Commun Phys 7, 235 (2024). https://doi.org/10.1038/s42005024017155
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DOI: https://doi.org/10.1038/s42005024017155
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