Abstract
The arrival time statistics of spin1/2 particles governed by Pauli’s equation, and defined by their Bohmian trajectories, show unexpected and very well articulated features. Comparison with other proposed statistics of arrival times that arise from either the usual (convective) quantum flux or from semiclassical considerations suggest testing the notable deviations in an arrival time experiment, thereby probing the predictive power of Bohmian trajectories. The suggested experiment, including the preparation of the wave functions, could be done with presentday experimental technology.
Introduction
In nonrelativistic quantum mechanics, the probability of finding a particle in a small spatial volume d^{3}r around position r at a fixed time t is given by Born’s rule Ψ(r, t)^{2}d^{3}r, where Ψ(r, t) is the wave function of the particle. This formula is experimentally well established. However, a formula for the probability of finding the particle at a fixed point r between times t and t + dt is the matter of an ongoing debate^{1,2,3,4,5,6}. Let us consider a typical time of arrival experiment, in which a particle is initially trapped in a region \({\rm{\Sigma }}\,\subset \,{{\mathbb{R}}}^{3}\), e.g., the interior of a potential well. The trap is released at, say, t = 0, allowing the particle to propagate freely in space. If the trapping potential is deep enough, the wave function of the particle at this instant, Ψ_{0}(r) ≡ Ψ(r, t = 0), practically vanishes outside the region Σ. Particle detectors placed on the boundary ∂Σ measure the time of arrival of the particle, denoted by τ. If the experiment is repeated many times, the recorded arrival times are random, even if the initial wave function of the particle (Ψ_{0}) is kept unchanged in each experiment. What is the probability distribution of arrival times \({{\rm{\Pi }}}^{{{\rm{\Psi }}}_{0}}(\tau )\) as a functional depending on Ψ_{0} and on ∂Σ?
Measurement in quantum mechanics has become in recent decades a tricky notion. Traditionally, measurement outcomes were solely associated with observables, represented by selfadjoint operators on the Hilbert space of the measured system. However, it has long been known that for time measurements, such as arrival times, no such observable exists [2, §8.5]. In fact, the notion of selfadjoint operators defining quantum observables does not apply to many other experiments as well. To remedy this situation, the notion of an observable was generalized to positive operator valued measures (POVMs), and there are various suggestions for arrival time POVMs (see^{7,8} for a discussion). In Dürr et al.^{9}, all measurements describable by POVMs were called linear measurements. Another class of measurements have also been performed, the socalled weak measurements^{10}, which are nonlinear in the sense of^{9}. So far, however, no theoretical predictions for the arrival time distribution \({{\rm{\Pi }}}^{{{\rm{\Psi }}}_{0}}(\tau )\) have been backed up by experiments (see^{1,2,8,11,12,13} for various proposals). On the other hand, recent ‘attoclock’ experiments (claimed to be measuring the tunnel delay time of ionized electrons) have shown some of the theoretical ideas to be empirically inadequate^{14,15}.
One problem with arrival time measurements is that detection events are based on interactions of the detectors with the detected particle, which may disturb its wave function in an uncontrollable way, leading to backscattering and in extreme cases to the quantum Zeno effect. While this is a valid concern, we note that the doubleslit experiment (mentioned also below) is an example where the distribution of arrival positions of the particle on the detector screen (the ubiquitous interference picture) is analyzed without any reference to the presence of the detector. Note well that the particles strike the detector surface at random times, a fact blissfully ignored in the usual discussions of the doubleslit experiment–and there are good reasons why that is justified. We expect that in the experiment proposed in this paper the same will be true, i.e., the detection event should not be drastically disturbed by the presence of the detector. Our expectation is based on our results, namely, on the very striking articulated features of the computed arrival time distributions, which should survive mild disturbances.
A further problem is that the notion of arrival time is most naturally connected with that of particle trajectories, an idea which is hard to concretize in the orthodox interpretation of quantum mechanics. Bohmian mechanics (or de BroglieBohm pilot wave theory) is a quantum theory (and not simply an alternative interpretation of quantum mechanics) where particles move on well defined smooth trajectories, hence it is naturally suited for computing arrival times of a particle. See Kocsis et al.^{10} for a weak measurement of average quantum trajectories in a doubleslit experiment, which can indeed be seen as Bohmian trajectories. Bohmian mechanics has been proven to be empirically equivalent to standard quantum mechanics, wherever the latter is unambiguous (e.g., in position and momentum measurements)^{9,16}. It has been shown that Bohmian mechanics provides in (far field) scattering situations ideal (i.e. devise independent) arrival time statistics for spin0 particles^{6}, via the quantum flux (or the probability current) J:
Only in scattering situations (i.e., when the detector surface ∂Σ is far away from the support of the initial wave function Ψ_{0}), is the surface integral in (1) demonstrably positive^{6}, otherwise it cannot be interpreted as a probability density (see^{7} for a discussion of POVMs versus flux statistics). Although often not recognized or emphasized in textbooks, it is the quantum flux J, integrated over time, that yields the doubleslit interference pattern of arrival positions of particles on the screen—as there is no given time at which the particles arrive at the screen.
Here, we must mention as well another flux based (Bohmian) arrival time distribution derived by C. R. Leavens (Eq. (9) of^{17}), which is valid only in one space dimension (as discussed in^{1}, [Ch. 5]). Therefore, it is not applicable for particles with spin1/2, except perhaps in some idealized situations. In this paper, we propose a Bohmian formula for the distribution of first arrival times of a spin1/2 particle (Eq. (9) below), which may be referred to as an ideal or intrinsic distribution, since it is formulated without referring to any particular measurement device, just as equation (1). In fact, the spin1/2 analogue of (1) becomes a special case of (9), whenever the socalled current positivity condition^{6} is met. Evaluating our formula numerically for a specific, carefully chosen experiment, we find that the resulting arrival time distributions show drastic and unexpected changes when control parameters are varied. Since the predicted distributions show such interesting and significant behavior, we suggest that the proposed experiment be performed to test the predictive power of Bohmian mechanics for spin1/2 particles.
One may legitimately ask, how can the idea of an ideal first arrival time distribution be entertained at all? Our standpoint on this is as follows: Given the relative simplicity of making the Bohmian prediction for the ideal arrival time distribution, and given the ambiguity of quantum mechanical proposals mentioned above, why not just do the experiment to check it? However the experimental results turn out, they provide in any case valuable experimental data that would enrich our understanding of quantum mechanics and of Bohmian mechanics as well. A word of clarification is, however, in order: we do not construe in our work any contradiction to quantum mechanics, since an unambiguous answer to the question of arrival times has not been given within this framework. The main advantage of Bohmian mechanics is the clear picture of reality, independent of observation, it provides and which in the problem of arrival times allows an unambiguous answer in contrast to the answers given in quantum mechanics, so far.
We proceed now to a description of the proposed experiment. A spin1/2 particle of mass m is constrained to move within a semiinfinite cylindrical waveguide (Fig. 1). Initially, it is trapped between the end face of the waveguide and an impenetrable potential barrier placed at a distance d. At the start of the experiment, the particle is prepared in a ground state Ψ_{0} of this cylindrical box, then the barrier at d is suddenly switched off, allowing the particle to propagate freely within the waveguide. The arrival surface ∂Σ is the plane situated at distance L (>d) from the end face of the waveguide. We then compute numerically from the Bohmian equations of motion how long it takes for the particle to arrive at ∂Σ, and determine the empirical distribution \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{{{\rm{\Psi }}}_{0}}(\tau )\) from typical trajectories (trajectories whose initial points are randomly drawn from the Born Ψ_{0}^{2} distribution) for different initial ground state wave functions.
The Bohmian equations for spin1/2 particles are as follows: The wave function Ψ(r, t) is a twocomponent complexvalued spinor solution of the Pauli equation
with given initial condition Ψ_{0}(r). Here, V(r, t) is an external potential, and \({\boldsymbol{\sigma }}\,=\,{\sigma }_{{\rm{x}}}\,\hat{{\boldsymbol{x}}}+{\sigma }_{{\rm{y}}}\,\hat{{\boldsymbol{y}}}+{\sigma }_{{\rm{z}}}\,\hat{{\boldsymbol{z}}}\) is a 3vector of Pauli spin matrices. The quantum continuity equation for the Pauli equation reads^{18,19}
where \({{\rm{\Psi }}}^{\dagger }\) is the adjoint of Ψ, and \({\rm{\Psi }}{}^{2}={{\rm{\Psi }}}^{\dagger }{\rm{\Psi }}\). The rightmost equality of (3) defines the Bohmian velocity field \({{\boldsymbol{v}}}_{\mathrm{Bohm}}^{{\rm{\Psi }}}={{\bf{J}}}_{\mathrm{Pauli}}/{\rm{\Psi }}{}^{2}\). The term \(\frac{\hslash }{m}{\rm{I}}{\rm{m}}[{{\rm{\Psi }}}^{\dagger }{\boldsymbol{\nabla }}{\rm{\Psi }}]\) is the socalled convective flux, while \(\frac{\hslash }{2m}\,{\boldsymbol{\nabla }}\times ({{\rm{\Psi }}}^{\dagger }{\boldsymbol{\sigma }}{\rm{\Psi }})\) is the spin flux. The alert reader will recognize that the spin flux is divergenceless, hence one may argue that neither the flux J_{Pauli} nor the Bohmian velocity are uniquely defined. However, observing that the Pauli equation and its flux emerge as nonrelativistic limits of the Dirac equation and the Dirac flux, respectively, which are unique^{20,21,22}, one is led directly to the current and Bohmian velocity given here.
The Bohmian trajectories are integral curves of the velocity field \({{\boldsymbol{v}}}_{\mathrm{Bohm}}^{{\rm{\Psi }}}\), hence the Bohmian guidance law reads
Here, R(t) is the position of the particle at time t. In Bohm’s theory, the spin1/2 particle has no degrees of freedom other than those specifying its position in space, so spin is not an extra degree of freedom. Thus, all quantum mechanical phenomena attributed to spin (such as the deflection of particles in the SternGerlach experiment) arise solely from the nonlinear equation of motion^{23,24}. The guidance law ((4) is time reversal invariant and its righthand side transforms as a velocity under Galilean transformations (see^{16,21} for further discussion). We integrate Eq. (4) for a statistical ensemble of Ψ_{0}^{2}distributed initial particle positions R(0) (see^{25} for a justification).
Methods
Our computations and results are for the following setup: Let the cylindrical waveguide be mounted on the xy−plane of a righthanded orthogonal coordinate system, the axis of the cylinder defining the z−axis (Fig. 1). Employing cylindrical coordinates r ≡ (ρ, ϕ, z), we model the potential field of the waveguide as \(V({\boldsymbol{r}},t)={V}_{\perp }(\rho )+{V}_{\parallel }(z,t)\), where \({V}_{\perp }(\rho )=\frac{1}{2}\,m\,{\omega }^{2}{\rho }^{2}\) is a transverse confining potential, and \({V}_{\parallel }(z,t)=v(z)+\theta (t)v(dz)\) is a time dependent axial potential comprised of an impenetrable hardwall at z = 0, viz.,
and another impenetrable potential barrier v(d−z) that is switched off at t = 0. The wall at z = 0 is the end face of the waveguide and θ(x) is Heaviside’s step function. Fortunately, near perfect harmonic confinements can be realized in conventional (ultrahigh vacuum) Penning traps, which can trap single electrons^{26,27} and protons^{28} over a wide range of trapping frequencies. For electrons, typical waveguide parameters read: L ≈ 6−10 mm, ω ≈ 10^{9}−10^{11} rad/s^{26,29}. In^{30} we give a detailed analysis of the proposed experiment for a quadrupole ion trap (Paul trap) waveguide.
The particle is prepared in a ground state of the cylindrical box at t = 0, which can be written as Ψ_{0}(r) = ψ_{0}(r)χ, where (setting ħ = m = d = 1),
is the spatial part of the wave function, and
is a normalized Bloch spinor \({(\chi }^{\dagger }\chi =1)\). Fixing α and β, we obtain different ground state wave functions. For instance, α = 0 (π) gives the spinup (spindown) ground state wave function, usually denoted by Ψ_{↑} (Ψ_{↓}), while \(\alpha \,=\,\frac{\pi }{2}\) yields the socalled updown ground state wave function \({{\rm{\Psi }}}_{\updownarrow }\,=\,\frac{1}{\sqrt{2}}({{\rm{\Psi }}}_{\uparrow }+{{\rm{\Psi }}}_{\downarrow })\). We refer to α and β as spin orientation angles, because they specify the orientation of the “spin vector” \({\bf{s}}\,:\,=\,\frac{1}{2}({{\rm{\Psi }}}_{0}^{\dagger }{\boldsymbol{\sigma }}{{\rm{\Psi }}}_{0})/{{\rm{\Psi }}}_{0}{}^{2}\), given by
The instant the barrier is switched off, the wave function spreads dispersively, filling the volume of the waveguide. The particle moves according to (4) on the Bohmian trajectory \({\boldsymbol{R}}(t)=R(t)\,[\,\cos \,{\rm{\Phi }}\,(t)\,\hat{{\boldsymbol{x}}}\,+\,\sin \,{\rm{\Phi }}\,(t)\,\hat{{\boldsymbol{y}}}\,]+Z(t)\hat{{\boldsymbol{z}}}\). In this choice of coordinates the first arrival time of a trajectory starting at R(0) and arriving at z = L is
where Z(t, R(0)) ≡ Z(t) is the z−coordinate of the particle at time t, and supp(Ψ_{0}) denotes the support of the initial wave function (the interior of the cylindrical box). Since the initial position R(0) is Ψ_{0}^{2}distributed^{25}, the distribution of τ(R(0)) is given by
In general, there is no closedform expression for (8), hence the integral in (9) cannot be evaluated analytically. However, if the Bohmian trajectories cross ∂Σ at most once, or in other words if the quantum flux J_{Pauli} is outward directed at every point of ∂Σ, at all times (also referred to as the current positivity condition), then \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{{{\rm{\Psi }}}_{0}}\,(\tau )\) reduces to the integrated quantum flux Π_{qf}(τ), Eq. (1) with the Pauli current J_{Pauli} replacing J^{5,6}.
Note well that by the very meaning of the quantum flux, (1) is a natural guess for the arrival time distribution from the point of view of standard quantum mechanics as well^{5}. However, (1) makes sense only if the lefthand side is positive, which need not be the case. Of course, if the current positivity condition holds, the lefthand side of (1) ≥0, and Π_{qf} becomes a special case of (9). Generally, this condition does not hold, in which case one computes \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{{{\rm{\Psi }}}_{0}}(\tau )\) numerically from a large number of Bohmian trajectories.
Results and Discussion
(i) For the spinup (α = 0) and spindown (α = π) wave functions the arrival time distribution \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{{{\rm{\Psi }}}_{0}}\,(\tau )\) coincides with the quantum flux expression (1), since in these cases the current positivity condition is satisfied. Moreover, in these cases J_{Pauli}(r,τ) can be replaced by the convective flux \(\frac{\hslash }{m}{\rm{I}}{\rm{m}}[{{\rm{\Psi }}}^{\dagger }{\boldsymbol{\nabla }}{\rm{\Psi }}]\) in (1). The resulting distribution has a heavy tail \(\sim {\tau }^{4}\) as τ → ∞. (ii) For other initial wave functions \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{{{\rm{\Psi }}}_{0}}\,(\tau )\) differs from (1) and falls off faster than τ^{−4}. For any initial ground state wave function, the arrival time distribution displays an infinite sequence of selfsimilar lobes below \(\tau \,=\,\frac{mdL}{2\pi \hslash }\) (see Fig. 2 below), which diminish in size as τ → 0. These lobes mirror typical wave function evolution when suddenly released to spread freely into the volume of the waveguide (see also^{31}). (iii) If the initial wave function is an equal superposition of the spinup and spindown wave functions (\(\alpha \,=\,\frac{\pi }{2}\)), the arrival time distribution pinches off at a maximum arrival time τ_{max}, i.e., no particle arrivals occur for τ > τ_{max}. Moreover an even more striking manifestation of the lobes can be seen: characteristic “noarrival windows” appear between the smaller lobes, inside which the arrival time distribution is zero. (iv) Time of flight measurements refer in general to semiclassical expressions based on the momentum distribution. Our distributions deviate significantly from this alleged semiclassical formula.
A few details concerning the computation of our results are in order. First note that the Pauli equation (2) with initial condition Ψ_{0}(r) can be solved in closed form, facilitating very fast numerical computation of Bohmian trajectories. The time dependent wave function takes the form Ψ(r, t) = ψ(r, t)χ, where the spin part is given by (7), while
where
which we call the ‘time evolution integral’. In (11):
where erfc(x) is the complementary error function. A detailed derivation of this result will be given elsewhere^{32}. Substituting our solution for the time dependent wave function in the guidance law (4), we obtain coupled nonlinear equations of motion for the spin1/2 particle:
where R, Φ and Z are the cylindrical coordinates of R and W′ = ∂W/∂z. The parameters in the initial wave function are, in view of (7), α and β, so we denote \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{{{\rm{\Psi }}}_{0}}\,(\tau )\,\equiv \,{{\rm{\Pi }}}_{\mathrm{Bohm}}^{\alpha \beta }\,(\tau )\). In fact, it is enough to consider β = 0 only since
These properties follow partly from the symmetries of Eq. (13) and partly from the initial uniform distribution of Φ(0) due to the choice of the initial state Eq. (6) (see^{32} for further details). We sample N ≈ 10^{5} initial positions from the Ψ_{0}^{2} distribution, solve Eq. (13) numerically for each point in this ensemble, continuing until the trajectory hits z = L, then record the arrival time and plot the histogram for \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{\alpha \mathrm{0}}(\tau )\).
For the spinup and spindown wave functions, Eq. (13c) reduces to \(\dot{Z}\,=\,{\rm{Im}}[W^{\prime} /W](Z(t),t)\) and numerically it turns out that Im[W′/W](L, t) > 0. Hence the spinup and spindown trajectories cross ∂Σ at most once, and the first arrival time distribution (or simply the arrival time distribution) in these cases equals (cf. Eq. (1))
As noted, (15) is nonnegative for all values of τ, and features prominently a large main lobe for \(\tau > \frac{L}{2\pi }\). The main lobe falls off as \({(\frac{L}{\pi })}^{3}\frac{4}{{\tau }^{4}},\,{\rm{a}}{\rm{s}}\,\,\tau \to {\rm{\infty }}.\) An infinite train of smaller lobes permeates the interval \(0 < \tau < \frac{L}{2\pi }\), which are well approximated by the formula \(\frac{4\pi }{L}\sin {c}^{2}(\frac{L}{\tau })\), whenever \(\tau \ll L\). Apart from that we find that Π_{qf}(τ) is a function only of the arrival distance L. It is also independent of the trapping frequency ω, which is rather surprising. In fact, the spinup (down) arrival time distributions are independent of the exact shape of the transverse confining potential V_{⊥}(ρ) of the waveguide as well^{32}. Figure 2 depicts our results for L = 100 (≈5 mm for a d = 50 m trap) and ω = 10^{3} (≈46.3 × 10^{6} rad/s). Note: We have expressed L, ω and τ in units of d, \(\frac{\hslash }{m{d}^{2}}\) and \(\frac{m{d}^{2}}{\hslash }\), respectively. For a d = 50 μm trap, known electron mass m ≈ 9.11 × 10^{−31} kg and reduced Planck’s constant ħ ≈ 1.05 × 10^{−34}Js, the frequency and time units are ≈46.3 × 10^{3} rad/s and ≈21.7 μs, respectively.
For wave functions corresponding to \(0 < \alpha \le \frac{\pi }{2}\) (cf. (14b)), the first arrival time distribution is not given by the integrated flux (1). This is because the Bohmian trajectories in these cases cross ∂Σ more than once, hence the aforementioned current positivity condition is not met. As α approaches \(\frac{\pi }{2}\), the tail of \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{\alpha \mathrm{0}}(\tau )\) thins gradually, pinching off completely at a characteristic maximum arrival time τ_{max} for \(\alpha =\frac{\pi }{2}\) (i.e. all Bohmian trajectories with wave function \({{\rm{\Psi }}}_{\updownarrow }\) strike the detector surface z = L before t = τ_{max}). In Fig. 2 τ_{max} ≈ 42.9, which corresponds to ≈1 ms. This behavior results in a sharp drop in the mean first arrival time τ in the vicinity of \(\alpha =\frac{\pi }{2}\), as shown in Fig. 3 below.
Unlike the spinup (down) case, the updown arrival time statistics are influenced by the trapping frequency ω. Keeping L fixed, we find that the maximum arrival time τ_{max}, mean first arrival time 〈τ〉, and the standard deviation σ corresponding to \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{\frac{\pi }{2}\mathrm{0}}(\tau )\) decrease with increasing ω, each approaching a constant for \(\omega \gg 1\). Conversely, when these quantities are graphed as functions of L with ω fixed, we see a clear linear growth in each (see Fig. 4).
Remarkably, these effects persist even for large L, provided ω is also made suitably large (a welcome feature for Penning traps). Increasing L causes the arrival time distributions to shift to larger values of τ, thus the smaller lobes can be easily seen, especially in experiments incapable of resolving very small arrival times.
Since the smaller (selfsimilar) lobes become progressively smaller, resolving the n^{th} lobe (denoting the main lobe by n = 1) is limited by the least time count (δt) of the measuring apparatus. Roughly, δt < width of n^{th}lobe should suffice to resolve the first n lobes. This translates into^{32}
For d = 50 μm and L = 5 mm (Fig. 2), a modest δt ≈ 10 μs will successfully resolve 8 lobes (main + 7 smaller lobes), while δt ≈ 0.1 μs will resolve as many as 83 lobes (main + 82 smaller lobes). However, we must also understand that only a few data points (about \((\frac{2}{{\pi }^{2}})\frac{N}{{n}^{4}}\) in N experiments) contribute to the n^{th} lobe, especially when \(n\gg 1\). This number, being independent of any tunable parameters like L, ω, etc., sets an intrinsic limit on the experimenter’s ability to resolve the distant lobes.
Finally, we come to the semiclassical arrival time distribution (dashed line in Fig. 2),
routinely used in the interpretation of timeofflight experiments. Here, \({\mathop{\psi }\limits^{ \sim }}_{0}\) denotes the Fourier transform of the initial wave function ψ_{0}. Although (17) is based on the tacit assumption that the particle moves classically between preparation and measurement stages (see § 5.3.1 of^{5}), it can also be motivated from the scattering formalism^{6}, [pg 971], provided the detector surface (∂Σ) is placed far away from the support of the initial wave function ψ_{0} (far field regime), where the external potentials are negligible. In typical cold atom experiments these conditions are met, hence the semiclassical formula (17) is empirically adequate. Therefore, soliciting deviations from (17), theorists have recommended “moving the detectors closer to the region of coherent wave packet production, or closer to the interaction region”^{2,33}, pg 419] (i.e. L ≈ d). However, such a relocation may disturb the wave function of the particle in an undesirable way.
For a meaningful comparison with our results, Eq. (17) (which is only applicable for free propagation) must be generalized to account for the presence of the waveguide. This is done with the help of Newtonian trajectories for our setup and the quantum mechanical distribution of the initial momenta of the particle. A careful calculation^{32} yields
which falls off as \((\frac{8}{{\pi }^{3}})\frac{L}{{\tau }^{2}}\) (compare this with Eq. (15) and its τ^{−4} falloff). In Fig. 2, we see that the semiclassical formula (18) for L = 100 remains distinctly different from the Bohmian arrival time distributions, notwithstanding the largeness of L. This is because the (spin dependent) Bohmian trajectories for our setup do not approach Newtonian trajectories even at large L, unlike the Bohmian trajectories of a free particle, which become approximately Newtonian at large distances^{6}.
Bohmian mechanics has been fruitfully applied in various physical disciplines with reference to technical advances^{34,35,36,37}, and to a better understanding of quantum phenomena, even in quantum gravity^{38,39}. We have proposed here a simple experiment that can possibly be performed with a detection mechanism as discussed in^{13,40,41,42}, and we have provided the Bohmian arrival times as a benchmark: Our results demonstrate that the distribution of first arrival times of a spin1/2 particle bears clear signatures of Bohm’s guidance law. The deviations of \({{\rm{\Pi }}}_{\mathrm{Bohm}}^{{{\rm{\Psi }}}_{0}}(\tau )\) from the quantum flux distribution (1), so strikingly in evidence, are particularly noteworthy.
Data Availability
The datasets generated and analyzed during the current study are available from the corresponding author on reasonable request.
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Acknowledgements
The authors would like to thank J. M. Wilkes for critically reviewing the manuscript. He and Markus Nöth verified our analytical calculations meticulously, which led to the correction of a mistake in the early stages of the work. The assistance of Grzesio Gradziuk and Leopold Kellers with the numerical simulations was essential, and greatly appreciated. Thanks are also due to Serj Aristarkhov, Lukas Nickel, Ward Struyve, Nikolai Leopold, Roderich Tumulka and Nicola Vona for inspiration and enriching discussions on the subject of this letter.
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S.D. conceived the experiment and conducted the numerical computations. The paper was written equally by S.D. and D.D.
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Das, S., Dürr, D. Arrival Time Distributions of Spin1/2 Particles. Sci Rep 9, 2242 (2019). https://doi.org/10.1038/s41598018382614
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