Nitrogen-Vacancy Color Centers Created by Proton Implantation in a Diamond

We present an experimental study of the longitudinal and transverse relaxation of ensembles of negatively charged nitrogen-vacancy (NV−) centers in a diamond monocrystal prepared by 1.8 MeV proton implantation. The focused proton beam was used to introduce vacancies at a 20 µm depth layer. Applied doses were in the range of 1.5×1013 to 1.5×1017 ions/cm2. The samples were subsequently annealed in vacuum which resulted in a migration of vacancies and their association with the nitrogen present in the diamond matrix. The proton implantation technique proved versatile to control production of nitrogen-vacancy color centers in thin films.     

Thermal Properties of Bayfol® HX200 Photopolymer

Bayfol^® HX200 photopolymer is a holographic recording material used in a variety of applications such as a holographic combiner for a heads-up display and augmented reality, dispersive grating for spectrometers, and notch filters for Raman spectroscopy. For these systems, the thermal properties of the holographic material are extremely important to consider since temperature can affect the diffraction efficiency of the hologram as well as its spectral bandwidth and diffraction angle. These thermal variations are a consequence of the distance and geometry change of the diffraction Bragg planes recorded inside the material. Because temperatures can vary by a large margin in industrial applications (e.g., automotive industry standards require withstanding temperature up to 125°C), it is also essential to know at which temperature the material starts to be affected by permanent damage if the temperature is raised too high. Using thermogravimetric analysis, as well as spectral measurement on samples with and without hologram, we measured that the Bayfol^® HX200 material does not suffer from any permanent thermal degradation below 160°C. From that point, a further increase in temperature induces a decrease in transmission throughout the entire visible region of the spectrum, leading to a reduced transmission for an original 82% down to 27% (including Fresnel reflection). We measured the refractive index change over the temperature range from 24°C to 100°C. Linear interpolation give a slope 4.5×10−4K−1 for unexposed film, with the extrapolated refractive index at 0°C equal to n0=1.51. This refractive index change decreases to 3×10−4K−1 when the material is fully cured with UV light, with a 0°C refractive index equal to n0=1.495. Spectral properties of a reflection hologram recorded at 532 nm was measured from 23°C to 171°C. A consistent 10 nm spectral shift increase was observed for the diffraction peak wavelength when the temperature reaches 171°C. From these spectral measurements, we calculated a coefficient of thermal expansion (CTE) of 384×10−6K−1 by using the coupled wave theory in order to determine the increase of the Bragg plane spacing with temperature.     

From the Cover: Individual variations lead to universal and cross-species patterns of social behavior

The duration of interaction events in a society is a fundamental measure of its collective nature and potentially reflects variability in individual behavior. Here we performed a high-throughput measurement of trophallaxis and face-to-face event durations experienced by a colony of honeybees over their entire lifetimes. The interaction time distribution is heavy-tailed, as previously reported for human face-to-face interactions. We developed a theory of pair interactions that takes into account individual variability and predicts the scaling behavior for both bee and extant human datasets. The individual variability of worker honeybees was nonzero but less than that of humans, possibly reflecting their greater genetic relatedness. Our work shows how individual differences can lead to universal patterns of behavior that transcend species and specific mechanisms for social interactions.

How individuals in a community interact with each other gives rise to collective emergent properties of the community (1–5). It reflects the individuals’ personal preference, social roles, the external environment, and other numerous factors applicable to specific context. The distribution of interevent times, or waiting time between two consecutive events, for temporal social networks has been much studied because of its relation to information or disease spreading (3, 4). It has been shown that the heavy tail in the interevent time distribution is due to a decision-based queuing process, in which some tasks are more prioritized than others (6, 7). In contrast, the distribution of contact duration, instead of the interevent time, and its connection to the nature of social interactions have not been studied as much.We have measured the duration of interactions among thousands of honeybees (Apis mellifera) in a hive, well-known eusocial insects that are easy to experimentally manipulate. Among many possible modes of honeybee social interaction, we focused on trophallaxis, which is mouth-to-mouth liquid food transfer. Trophallaxis occurs not only for feeding but also for communication (8, 9), making it a model system to study social interactions and collective effects (10, 11). To measure the interaction time, all of the honeybees in a colony were fitted with a barcode (12). A high-resolution machine vision camera imaged them at the rate of one frame per second. Then a customized algorithm detected each interaction event by analyzing the images and identified each bee, its position, and its orientation (12) (Materials and Methods).Note that all of the honeybee data used in this work were originally generated by the authors for separate studies not for the purpose of testing our theory discussed in this paper or even to acquire the data needed for the theory. We used all of the data available to us, which were the trophallaxis social network data acquired in 2013 and analyzed in ref. 12 (1164_2013, 1140_2013, 1138_2013, 1174_2013, and 1170_2013, which are trials 1 to 5 in ref. 12) and the trophallaxis social network data acquired in 2016 from colonies with partial treatment with Juvenile Hormone analogue (13, 14) (789_JHA_2016 and 757_JHA_2016) (see Materials and Methods for more detail on the colony preparation). Our theory works for all of them, which indicates its robustness.Fig. 1A shows that the distribution of honeybee interaction duration is heavy tailed. The exponents of the power law are –2.4, –2.3, –2.2, –2.2, –2.2, –2.7, –2.0, and –2.0 for each dataset listed in the legend from the top (1164_2013) to the bottom (757_JHA_2016). If every bee were the same and every interaction happened by chance, one might naively expect to see a peaked distribution such as a Gaussian. However, the observed heavy-tailed distribution suggests heterogeneity or variability among the population.Open in a separate windowFig. 1.(A) Duration of trophallaxis as well as F2F events that are not detected as trophallaxis follows a heavy-tailed distribution. The first number in the dataset name represents the number of observed honeybees in each colony, while the last number represents the year the experiment was performed. Note that the number of individuals in the last two datasets is actually smaller than what was originally observed (676 and 639 instead of 789 and 757, respectively) because we analyzed the data after applying an additional filter to exclude the time series during which the colonies were perturbed by the treatment of JHA, but we decided to keep the name of the datasets. The numbers of interactions used to make the plot are 302,221, 205,787, 191,795, 259,923, 329,170, 1,207,778, 136,529, and 115,965 in the order the datasets are listed in the legend. (B) Human face-to-face (F2F) interaction in various settings exhibits a heavy-tailed distribution. The numbers of interactions used are 10,677, 19,774, 14,037, 32,425, 9,865, 77,520, 26,039, 4,591, and 33,750 in the order the datasets are listed in the legend. In both A and B, error bars indicate the SE. In B, lower error bars for bins with count 1 could not be drawn in logarithmic scale because it extends to 0.In order to improve the statistical power of our analysis, we also examined F2F events, where honeybees were close and oriented toward each other but not actually engaging in trophallaxis (15). F2F events occur about an order of magnitude more frequently than trophallaxis. F2F events include undetected trophallaxis and possible antennation, but the nature of the honeybee interaction during F2F is not as well defined as trophallaxis. Nevertheless, being F2F is a necessary but not sufficient condition for trophallaxis, so as long as the distance apart is not larger than the length of a honeybee, it would be expected that F2F events scale similarly to trophallaxis events. Indeed, coarse data for temporal networks retain some statistics of the actual interaction including heavy-tailedness of the contact duration distribution (16) (also confirmed by our results in Fig. 1A; see 1166_F2F_2013).In our work, the detection of trophallaxis events is estimated using a two-step filtering process because we are not able to observe directly fluid transfer through the proboscis. Our detection scheme is not subject to tracking error generated by bees that are misoriented or have other individual variations in visibility. Our method, described in detail in our earlier work (12), first selects bee pairs of close enough distance and orientation toward each other. Then we apply a second filter, selecting the bee pairs that are physically connected by proboscis through image processing. Through this second filter, we can filter out the pairs with inaccurate tracking of position or orientation because those pairs would not be connected by the proboscis. Even if there were innately poorly tracked bees, the multiple layers of filtering minimize the detection error on the trophallaxis events. There is quantitative evidence that there is minimal tracking error impact on our results, even without the second filter. The F2F dataset concerns bee pairs that satisfy only the geometric constraint regarding distance and orientation but are not detected as being connected by the proboscis. This dataset has not filtered out the pairs with inaccurate position and orientation through the second filter. However, as shown below, this F2F dataset gives the same power law and scaling laws as the trophallaxis datasets. In other words, the statistics of the data are retained regardless of whether the data have gone through the second filter, which is consistent with the notion that inaccuracy due to positional tracking error was already negligible after the first filter. Our earlier work (12) showed by subsampling of the trophallaxis interaction data at different sampling rates that the statistics of the trophallaxis interaction are robust against subsampling and false negatives, with the statistics of the times between trophallaxis events being robust to detection errors. This strongly suggests that the statistics of the duration of the trophallaxis events will also be robust.We compared the honeybee data with human data recorded by the SocioPatterns collaboration (16–24) to explore whether there are universal patterns of social interaction. Fig. 1B shows that human F2F interaction time in various settings also exhibits a heavy-tailed distribution. The exponents of the power law are –2.4, –2.7, –2.9, –3.4, –2.7, –3.6, –2.5, –2.9, and –2.6 for each dataset listed in the legend from the top (Highschool_2011) to the bottom (Workplace_2015). Such similarity across different systems indicates an unexpected universality governing the interaction in social systems and suggests that a minimal model (25) should be able to capture the salient features of the interactions.To construct such a minimal model, we treat the social bond between bees as an effective particle. We focus on bees here, but the model is also applicable to humans. The bond is the edge in the usual network representation of social interactions. This effective particle has two states representing an interacting pair and a noninteracting pair. The particle jumps from one state to the other with a fixed rate ω in the reaction coordinate, as depicted in Fig. 2. Although jumping happens in both directions, we focus on the jumping from the interacting state to the noninteracting state because the interaction time is the waiting time for the first jump in that direction. The distribution of the first jump time f(t,ω) is obtained by multiplying the probability not to jump at each time step until time t with the probability density to jump at time t. The first part, the probability not to jump until t, is (1−ωdt)t/dt, where dt is the interval for a time step. Taking the limit of infinitesimal time steps yields limdt→0(1−ωdt)t/dt=e−ωt. The distribution of the first jump time is then given by f(t,ω)=ωe−ωt.Open in a separate windowFig. 2.Schematic picture for the theory of interaction durations. The dotted circular area indicates the neighborhood of the center bee. The neighbors, or the potential partners, inside the circle are connected with the center bee by a bond. The bond between two bees is represented by a particle in the reaction coordinate. This particle has two states representing an interacting pair and a noninteracting pair, and it changes its state by jumping over the energy barrier E with rate ω. Among the neighbors, the center bee is assumed to primarily interact with a bee with the highest barrier because they form the stablest pair.Some pairs tend to have longer interactions than others, and this is reflected in their value of ω. To take into account this variability or heterogeneity within the community, we integrate the jump time distribution for a fixed ω over the distribution of rates p(ω). To determine p(ω), we use the Kramers theory for escape over a potential barrier (26): the distribution of ω is related to the distribution of energy barrier heights E through ω=ω0e−E where ω0 is a constant (26). Here we use a dimensionless energy scale E, which is already normalized with possible fluctuations.We propose that the barrier height distribution p(E) follows the extreme value distribution for maxima. As illustrated in Fig. 2, a given bee has multiple candidate partners with which to interact. Each possible pair is associated with a certain barrier height. The pair with the highest barrier would spend more time together because it is more difficult for the particle to escape. This partner is thus interpreted as the most likeable and ends up forming a pair. Then the observed energy barrier E is the maximum among the neighbors. The distribution of the maximum is taken to be the Fisher–Tippett–Gumbel distribution for maxima (27, 28), which is appropriate if the parent distribution for barrier heights is localized, as seems reasonable. Then we express p(E)=(α+1)e−(α+1)Ee−e−(α+1)E where α is an undetermined parameter. For large E, p(E)∼e−(α+1)E. We take the large E limit because the heavy-tailed behavior is observed at large t. Using ω=ω0e−E, we find that p(ω)∼ωα for small ω, equivalent to large E. Combining this p(ω) with the exponential pair interaction time distribution, we getf(t)=∫dωp(ω)f(ω,t)∼t−(2+α)(1)as the interaction time distribution for the population. The power law form suggests that the assumption about the parent distribution for barrier heights is valid. More details of this calculation are provided in SI Appendix. The quantity α in the exponent connects the community interaction time distribution f(t) with the distribution of barrier heights p(E).We remark that a similar derivation for the heavy-tailed time distribution as shown in Eq. 1 arises in the theory for defect jumping (29) and a model of traps (30, 31) in glass. However, the interpretation of p(E) in social interactions is different from the analogue in disordered materials. Atoms in a glass successively hop over multiple energy barriers in a rough potential landscape, so the integration over p(E) is an average over the energy barriers experienced by one atom. On the other hand, our particles for the bond between a pair jump over one energy barrier to change their state. Thus, the integration over p(E) is an ensemble average over the population.Next we turn to verifications of the predictions of this theory. The simple theory predicts an exponential pair interaction time distribution. The quantile–quantile plots for pair interaction times (Fig. S1) suggest that the pair interaction time distributions for both honeybees and humans are better expressed by hyperexponential distributions, which are weighted sums of two exponential distributions. The theory is not affected by this additive modification, as discussed in SI Appendix. There are so many pairs in each colony that it is not practical to show the goodness of hyperexponential fit for each pair separately. Therefore, we devised a data collapse to show the fit of all pairs in one figure. Only the pairs that yield more than seven points of evaluation for the empirical cumulative distribution function (ECDF) were considered. The cumulative distribution function (CDF) for a hyperexponential function is Y(t)=1−ge−ω1t−(1−g)e−ω2t, where g is the weight and ω1 and ω2 are the rates for each exponential. We define a new variable z≡ω1t and rewrite the CDF as F(z,g,ω1,ω2)≡(1/g)1−Y−(1−g)e−(ω2/ω1)z =e−z, where Y is ECDF. Then the x axis only depends on one variable z. If the data are well fitted by this functional form, plotting F(z,g,ω1,ω2) against e−z should produce a cloud of data points aligning with the y=x reference line.Fig. 3 A and B show that most honeybee and human pair interaction times are well fitted to hyperexponential distributions. Fig. 3 A and B, Insets, show the fit of a pair to provide some intuition of the fitting process. The fitted CDF tends to deviate more at small e−z, or large t, because the CDF value of the fitting function approaches 1 for t→∞, while the ECDF value is 1 at the longest observed interaction time.Open in a separate windowFig. 3.(A) Pair interaction time distributions for honeybee pairs are fitted to hyperexponential functions and collapsed together. The numbers of pairs used to generate the plot are 197, 99, 328, 443, 561, 1,806, 46, and 20 in the order the datasets are listed in the legend. The pair interaction time distributions of these pairs were fitted. (Inset) Fitting ECDF of pair interaction times of a pair from 1164_2013. (B) The same plot as A but with human pairs. The numbers of pairs used are 37, 82, 59, 58, 39, 143, 98, 15, and 171 in the order the datasets are listed in the legend. (Inset) Fitting ECDF of pair interaction times of a pair from Highschool_2011. (C) Comparison between the scaling of interaction time distribution and mean pair interaction time distribution for 1164_2013. (D) The same plot as C but for Primaryschool. The number of mean pair interaction times used is the same as the number of pairs used for fitting, which is listed in A and B. The number of interaction times used is the same as what is listed in Fig. 1. Error bars indicate the SE, and lower error bars for bins with count 1 could not be drawn in logarithmic scale because it extends to 0.A second prediction from the model is the exponential barrier height distribution. Although E is not a directly measurable variable, the relation ω=ω0e−E enables us to indirectly measure p(E) because the mean pair time associated with an energy barrier is τ=1/ω. The relation p(E)∼e−(α+1)E implies that p(τ)∼τ−(2+α) for large τ, which has the same exponent as f(t) in Eq. 1. Therefore, comparing the exponent of the tail of f(t) and p(τ) provides a test of the theory, in particular, the proposed functional form of p(E).Fig. 3 C and D demonstrate the same scaling between the tail of f(t) and p(τ) for 1164_2013 and Primaryschool, respectively. The comparison of scaling for seven other honeybee datasets and eight other human datasets is shown in SI Appendix, Fig. S4. Here τ is obtained from fitting parameters ω, not from averaging of pair interaction times, because τ is the mean pair interaction time associated with a single energy barrier (SI Appendix). If we retain the full form for p(E), i.e., including the superexponential term in the Fisher–Tippett–Gumbel distribution, p(τ) is expected to have a peak at small τ, which may explain the peak in Fig. 3C.One might think that the identical scaling between f(t) and p(τ) is a consequence of the so-called stable law (32) because τ is an average of t for pairs. Then, depending on the tail of f(t), the distribution of τ would be given either by the central limit theorem, i.e., a Gaussian, or by a power law with the same exponent as f(t). However, this is not correct, as explained in SI Appendix, because the parent distribution of p(τ) is not f(t) but instead is the pair interaction time distribution f(ω,t).The two energy barriers suggested by the hyperexponential pair interaction time distribution imply a multidimensional potential landscape of the reaction coordinate. Our model does not limit the number of barriers, allowing the pair interaction time distribution in principle to be an arbitrary sum ∑igiωie−ωit, but the weight of further barriers seems to be too small to contribute to the dynamics. It is evident that different pairs are characterized by different barrier heights, but whether it is a specific pair or a specific individual that determines the barrier height cannot be determined by the analysis so far.To explore the effect of individuality in social interactions, we calculated the Gini coefficient (33) for 1) the total interaction time spent by each individual, 2) the total number of interactions each individual had, and 3) the total number of partners with which each individual interacted. Widely used to express inequality in economics, the Gini coefficients have recently been used to quantify inequality in the activity level of eusocial insects (34, 35). Fig. 4 shows a graphical representation of the results, known as the Lorenz plot (36), for the total interaction time spent by bees and humans. The Lorenz plots for other variables are shown in SI Appendix, Fig. S5.Open in a separate windowFig. 4.Lorenz plots of the total time spent for interaction by honeybees and humans. (A) Gini coefficients of bees are as follows: 1164_2013, 0.2373; 1140_2013, 0.2111; 1138_2013, 0.3013; 1174_2013, 0.2760; 1170_2013, 0.2698; 1166_F2F_2013, 0.2089; 789_JHA_2016, 0.1941; and 757_JHA_2016, 0.1727. The numbers of data points used in the plot are the same as the numbers of individuals in each dataset, which are 1,164, 1,140, 1,138, 1,174, 1,170, 1,166, 676, and 639. (B) Gini coefficients of humans are as follows: Highschool_2011, 0.4333; Highschool_2012, 0.4879; Hospital, 0.5488; Household, 0.5012; Hypertext, 0.4576; Primaryschool, 0.2799; SFHH, 0.4937; Workplace_2013, 0.4493; and Workplace_2015, 0.3753. The numbers of individuals used are 126, 180, 75, 75, 113, 242, 403, 92, and 217.To read these results, note that in a Lorenz plot the greater the deviation from the y=x reference line, the closer the Gini coefficient to unity, thus indicating a greater level of inequality. More inequality in our data means a greater contribution by individuals to the dynamics, signifying the effect of individuality. Fig. 4A shows that the Gini coefficients for honeybees are in the range 0.2 to 0.3, whereas for humans they are in the range 0.3 to 0.5. Thus, although individual bees are distinct, they are not as different from each other as humans are (Fig. 4B). The Lorenz plots and Gini coefficients for the total number of interactions and total number of partners provided in Fig. S5 show the same trend. The reduced individuality in honeybees compared to humans might be due to the average coefficient of relatedness being r=0.75 among workers of the same colony as the queen was inseminated with a single male in these experiments, but further study is needed to verify this conjecture. Since the interaction time is a shared value between a pair, it is nontrivial to completely separate the contribution of individuals. The effect of individuality in social interactions is therefore an open question but one that we have provided the tools to explore. Nevertheless, along with earlier studies of possible chemosensory mechanisms for individual identification (37), our results provide confirmation and quantification of the conjecture from recent work on the personality of honeybee workers as described in ref. 38 that some individuals are more likely to be interactive and engaged in food sharing, while others are less so.The recently discovered heterogeneous food distribution in the Camponotus sanctus ant colony (39) may suggest individual variations in workers of this other well-known eusocial insect. Although the ratio of transferred food volume to maximal transferable volume during trophallaxis when the donor is a forager is measured to follow the same exponential distribution with the same parameter as the case when the donor is a nonforager (39), it does not necessarily mean the lack of individuality in ants because the individual variations may have been averaged out as the data of many pairs were analyzed collectively. If the data were analyzed for each pair, individual variations may have been observed. It is not the scope of our work, but it would be possible to study the effect of individuality on the food mixing due to trophallaxis of eusocial insects.We have shown that high-resolution tracking can yield detailed multiscale information about the interactions and behavior of individuals within a community. Our results suggest that individual differences can lead to patterns of behavior that are universal and transcend species, context, and specific mechanisms for social interactions.     

A backward-spinning star with two coplanar planets

It is widely assumed that a star and its protoplanetary disk are initially aligned, with the stellar equator parallel to the disk plane. When observations reveal a misalignment between stellar rotation and the orbital motion of a planet, the usual interpretation is that the initial alignment was upset by gravitational perturbations that took place after planet formation. Most of the previously known misalignments involve isolated hot Jupiters, for which planet–planet scattering or secular effects from a wider-orbiting planet are the leading explanations. In theory, star/disk misalignments can result from turbulence during star formation or the gravitational torque of a wide-orbiting companion star, but no definite examples of this scenario are known. An ideal example would combine a coplanar system of multiple planets—ruling out planet–planet scattering or other disruptive postformation events—with a backward-rotating star, a condition that is easier to obtain from a primordial misalignment than from postformation perturbations. There are two previously known examples of a misaligned star in a coplanar multiplanet system, but in neither case has a suitable companion star been identified, nor is the stellar rotation known to be retrograde. Here, we show that the star K2-290 A is tilted by 124○±6○ compared with the orbits of both of its known planets and has a wide-orbiting stellar companion that is capable of having tilted the protoplanetary disk. The system provides the clearest demonstration that stars and protoplanetary disks can become grossly misaligned due to the gravitational torque from a neighboring star.

The K2-290 system (1) consists of three stars. The primary star, K2-290 A, is a late-type F star with a mass of 1.19±0.07 solar masses (M⊙) and a radius of 1.51±0.07 solar radii (R⊙). The secondary star, K2-290 B, is an M dwarf with a projected orbital separation of 113±2 astronomical units (au). The tertiary star, K2-290 C, is another M dwarf located farther away, with a projected separation of 2467−155+177 au. The primary star harbors two transiting planets. The inner planet “b” has an orbital period of 9.2 d and a radius of 3.06±0.16 Earth radii (R⊕), making it a “hot sub-Neptune.” The outer planet “c” is a “warm Jupiter” with orbital period 48.4 d, radius 11.3±0.6 R⊕, and mass 246±15 Earth masses (M⊕).     

Study of Radiation Characteristics of Intrinsic Josephson Junction Terahertz Emitters with Different Thickness of Bi2Sr2CaCu2O8+δ Crystals

The radiation intensity from the intrinsic Josephson junction high-Tc superconductor Bi2Sr2CaCu2O8+δ terahertz emitters (Bi2212-THz emitters) is one of the most important characteristics for application uses of the device. In principle, it would be expected to be improved with increasing the number of intrinsic Josephson junctions N in the emitters. In order to further improve the device characteristics, we have developed a stand alone type of mesa structures (SAMs) of Bi2212 crystals. Here, we understood the radiation characteristics of our SAMs more deeply, after we studied the radiation characteristics from three SAMs (S1, S2, and S3) with different thicknesses. Comparing radiation characteristics of the SAMs in which the number of intrinsic Josephson junctions are N∼ 1300 (S1), 2300 (S2), and 3100 (S3), respectively, the radiation intensity, frequency as well as the characteristics of the device working bath temperature are well understood. The strongest radiation of the order of few tens of microwatt was observed from the thickest SAM of S3. We discussed this feature through the N2-relationship and the radiation efficiency of a patch antenna. The thinner SAM of S1 can generate higher radiation frequencies than the thicker one of S3 due to the difference of the applied voltage per junctions limited by the heat-removal performance of the device structures. The observed features in this study are worthwhile designing Bi2212-THz emitters with better emission characteristics for many applications.     

Isoelectronic perturbations to f-d-electron hybridization and the enhancement of hidden order in URu2Si2

Electrical resistivity measurements were performed on single crystals of URu_2–xOs_xSi₂ up to x = 0.28 under hydrostatic pressure up to P = 2 GPa. As the Os concentration, x, is increased, 1) the lattice expands, creating an effective negative chemical pressure P_ch(x); 2) the hidden-order (HO) phase is enhanced and the system is driven toward a large-moment antiferromagnetic (LMAFM) phase; and 3) less external pressure P_c is required to induce the HO→LMAFM phase transition. We compare the behavior of the T(x, P) phase boundary reported here for the URu_2-xOs_xSi₂ system with previous reports of enhanced HO in URu₂Si₂ upon tuning with P or similarly in URu_2–xFe_xSi₂ upon tuning with positive P_ch(x). It is noteworthy that pressure, Fe substitution, and Os substitution are the only known perturbations that enhance the HO phase and induce the first-order transition to the LMAFM phase in URu₂Si₂. We present a scenario in which the application of pressure or the isoelectronic substitution of Fe and Os ions for Ru results in an increase in the hybridization of the U-5f-electron and transition metal d-electron states which leads to electronic instability in the paramagnetic phase and the concurrent formation of HO (and LMAFM) in URu₂Si₂. Calculations in the tight-binding approximation are included to determine the strength of hybridization between the U-5f-electron states and the d-electron states of Ru and its isoelectronic Fe and Os substituents in URu₂Si₂.

The heavy-fermion superconducting compound URu2Si2 is known for its second-order phase transition into the so-called “hidden-order” (HO) phase at a transition temperature T0 ≈17.5 K. Extensive investigation of the phase space in proximity to the HO phase transition has provided a detailed picture of the electronic and magnetic structure of this unique phase (1–42). However, more than three decades after the initial characterization of URu2Si2 (1–3), the order parameter for the HO phase is still unidentified.Most perturbations to the URu2Si2 compound have the effect of suppressing HO. The application of an external magnetic field (H) suppresses the HO phase (41, 43) and many of the chemical substitutions (x) at the U, Ru, or Si sites that have been explored significantly reduce T0, even at modest levels of substituent concentration (44–52). At present, only three perturbations are known to consistently enhance the HO phase in URu2Si2: 1) external pressure P, 2) isoelectronic substitution of Fe ions for Ru, and 3) isoelectronic substitution of Os ions for Ru. Upon applying pressure P, the HO phase in pure URu2Si2 is enhanced (6) and the system is driven toward a large-moment antiferromagnetic (LMAFM) phase (53). The HO→LMAFM phase transition is identified indirectly by a characteristic “kink” at a critical pressure Pc ≈1.5 GPa in the T0 (P) phase boundary (18, 53, 54) and also directly by neutron diffraction experiments, which reveal an increase in the magnetic moment from μ ∼ (0.03±0.02)μB/U in the HO phase to μ ∼ 0.4 μB/U in the LMAFM phase (13, 55, 56).Recent reports indicate that the isoelectronic substitution of Fe ions for Ru in URu2Si2 replicates the T0(P) behavior in URu2Si2 (57–59). An increase in x in URu2−xFexSi2 enhances the HO phase and drives the system toward the HO→LMAFM phase transition at a critical Fe concentration xc ≈0.15 (58, 60). The decrease in the volume of the unit cell due to substitution of smaller Fe ions for Ru may be interpreted as a chemical pressure, Pch, where the Fe concentration x can be converted to Pch (x) (57, 59). In addition, the induced HO→LMAFM phase transition in URu2−xFexSi2 occurs at combinations of x and P that consistently obey the additive relationship: Pch(x) + Pc ≈1.5 GPa (57, 59). These results have led to the suggestion that Pch is equivalent to P in affecting the HO and LMAFM phases (58, 59).Reports of the isoelectronic substitution of larger Os ions for Ru have shown that an increase in x in URu2−xOsxSi2 1) expands the volume of the unit cell, thus creating an effective negative chemical pressure (Pch ≤0); 2) enhances the HO phase; and 3) drives the system toward a similar HO→LMAFM phase transition at a critical Os concentration of xc ≈0.065 (60–62). These results are contrary to the expectation that a negative Pch would lead to a suppression of HO and complicate the view of chemical pressure as a mechanism affecting the evolution of phases in URu2Si2.In this paper, we report on the behavior of the T(x, P) boundary for the URu2−xOsxSi2 system based on ρ(T) measurements of single crystals of URu2−xOsxSi2 as a function of Os concentration x and applied pressure P. The T(x, P) phase boundary observed here for the URu2−xOsxSi2 system (57–59) is compared to that of the URu2−xFexSi2 system and also with the behavior of T(P) in pure URu2Si2. As an explanation for the enhancement of HO toward the HO→LMAFM phase transition, we suggest a scenario in which each of the perturbations of Os substitution, Fe substitution, and pressure P favors delocalization of the 5f electrons and increases the hybridization of the uranium 5f-electron and transition metal (Fe, Ru, Os) d-electron states. To avoid an ad hoc explanation of the effect of increasing the Os concentration x in URu2−xOsxSi2, compared to the effects of pressure P and Fe substitution, we explain how pressure P, Fe substitution, and Os substitution are three perturbative routes to enhancement of the U-5f- and d-electron hybridization. The importance of the 5f- and d-electron hybridization to the emergence of HO/LMAFM is presented in the context of the Fermi surface (FS) instability that leads to a reconstruction and partial gapping of the FS during the transition from the paramagnetic (PM) phase to the HO and LMAFM phases (2, 6, 20, 22, 24–26, 37, 38, 63).In an effort to further understand the effect of isoelectronic substitution on the 5f- and d-electron hybridization, calculations in the tight-binding approximation were made for compounds from the series UM2Si2 (M = Fe, Ru, and Os). The calculations indicate that the degree of hybridization is largely dependent on the magnitude of the difference between the binding energy of the localized U-5f electrons and that of the transition metal d electrons.     

Mid-Infrared Laser Generation of Zn1−xMnxSe and Zn1−xMgxSe (x ≈ 0.3) Single Crystals Co-Doped by Cr2+ and Fe2+ Ions—Comparison of Different Excitation Wavelengths

Two different mid-infrared (mid-IR) solid-state crystalline laser active media of Cr2+, Fe2+:Zn1−xMnxSe and Cr2+, Fe2+:Zn1−xMgxSe with similar amounts of manganese or magnesium ions of x ≈ 0.3 were investigated at cryogenic temperatures for three different excitation wavelengths: Q-switched Er:YLF laser at the wavelength of 1.73 μm, Q-switched Er:YAG laser at 2.94 μm, and the gain-switched Fe:ZnSe laser operated at a liquid nitrogen temperature of 78 K at ∼4.05 μm. The temperature dependence of spectral and laser characteristics was measured. Depending on the excitation wavelength and the selected output coupler, both laser systems were able to generate radiation by Cr2+ or by Fe2+ ions under direct excitation or indirectly by the Cr2+→ Fe2+ energy transfer mechanism. Laser generation of Fe2+ ions in Cr2+, Fe2+:Zn1−xMnxSe and Cr2+, Fe2+:Zn1−xMgxSe (x ≈ 0.3) crystals at the wavelengths of ∼4.4 and ∼4.8 μm at a temperature of 78 K was achieved, respectively. The excitation of Fe2+ ions in both samples by direct 2.94 μm as well as ∼4.05 μm radiation or indirectly via the Cr2+→ Fe2+ ions’ energy transfer-based mechanism by 1.73 μm radiation was demonstrated. Based on the obtained results, the possibility of developing novel coherent laser systems in mid-IR regions (∼2.3–2.5 and ∼4.4–4.9 μm) based on AIIBVI matrices was presented.     

Cutting Tests of the Outer Layer of Material Using Onion as an Example

This paper describes experimental research on cutting the outer layer of onions in the machine peeling process. The authors’ own globally innovative modular machine construction was used for this purpose. The onion peeling machine was constructed on a real scale. The effectiveness of the machine’s functioning Se was defined as the ratio of the mass of material correctly removed by the scale blower mp to the mass of all material leaving the machine on the test bench mc. In order to carry out the experimental research, a test stand was constructed, a research plan and programme were adopted, and the research methodology was developed. The results obtained during the experimental research and the data obtained from the regression function equations for the developed design of the onion peeling machine were used to build systems of independent variables, for which the dependent variable Se reached extreme values. The effectiveness of the machine’s operation Se of modular construction increased with the increase in the depth of the external incisions of the shells dn, the number of scale-blowing nozzles, and the pressure of the air supply to the scale-blowing unit p. Increasing the material feed rate vp and the distance of the air nozzles from the material to be processed hd reduced the machine’s efficiency Se. The tests carried out showed a high level of efficiency on the level of Se=0.645−0.780, which is not found in mass-produced machines.     

Faraday Rotation of Dy2O3, CeF3 and Y3Fe5O12 at the Mid-Infrared Wavelengths

The relatively narrow choice of magneto-active materials that could be used to construct Faraday devices (such as rotators or isolators) for the mid-infrared wavelengths arguably represents a pressing issue that is currently limiting the development of the mid-infrared lasers. Furthermore, the knowledge of the magneto-optical properties of the yet-reported mid-infrared magneto-active materials is usually restricted to a single wavelength only. To address this issue, we have dedicated this work to a comprehensive investigation of the magneto-optical properties of both the emerging (Dy2O3 ceramics and CeF3 crystal) and established (Y3Fe5O12 crystal) mid-infrared magneto-active materials. A broadband radiation source was used in a combination with an advanced polarization-stepping method, enabling an in-depth analysis of the wavelength dependence of the investigated materials’ Faraday rotation. We were able to derive approximate models for the examined dependence, which, as we believe, may be conveniently used for designing the needed mid-infrared Faraday devices for lasers with the emission wavelengths in the 2-μm spectral region. In the case of Y3Fe5O12 crystal, the derived model may be used as a rough approximation of the material’s saturated Faraday rotation even beyond the 2-μm wavelengths.     

XPS Study in BiFeO3 Surface Modified by Argon Etching

This paper reports an XPS surface study of pure phase BiFeO3 thin film produced and later etched by pure argon ions. Analysis of high-resolution spectra from Fe 2p, Bi 4f and 5d, O 1s, and the valence band, exhibited mainly Fe3+ and Bi3+ components, but also reveal Fe2+. High-energy argon etching induces the growth of Fe(0) and Bi(0) and an increment of Fe2+, as expected. The BiFeO3 semiconductor character is preserved despite the oxygen loss, an interesting aspect for the study of the photovoltaic effect through oxygen vacancies in some ceramic films. The metal-oxygen bonds in O 1s spectra are related only to one binding energy contrary to the split from bismuth and iron reported in other works. All these data evidence that the low-pressure argon atmosphere is proved to be efficient to produce pure phase BiFeO3, even after argon etching.     

Doping evolution of the Mott–Hubbard landscape in infinite-layer nickelates

The recent observation of superconductivity in Nd0.8Sr0.2NiO2 has raised fundamental questions about the hierarchy of the underlying electronic structure. Calculations suggest that this system falls in the Mott–Hubbard regime, rather than the charge-transfer configuration of other nickel oxides and the superconducting cuprates. Here, we use state-of-the-art, locally resolved electron energy-loss spectroscopy to directly probe the Mott–Hubbard character of Nd1−xSrxNiO2. Upon doping, we observe emergent hybridization reminiscent of the Zhang–Rice singlet via the oxygen-projected states, modification of the Nd 5d states, and the systematic evolution of Ni 3d hybridization and filling. These experimental data provide direct evidence for the multiband electronic structure of the superconducting infinite-layer nickelates, particularly via the effects of hole doping on not only the oxygen but also nickel and rare-earth bands.

The discovery of copper oxide high-temperature superconductors (1) and subsequent realizations of non–copper-based compounds (2, 3) has spurred significant efforts not only to understand the mechanisms underlying superconductivity (4–7) but also to identify and realize new host materials. The recent discovery of superconducting nickel-based thin films (8) has opened an opportunity to explore a system that is closely related to the infinite-layer cuprates in both crystal structure and transition-metal electron count. Trading out the Cu2+ ions for Ni1+ preserves an isoelectronic formal 3d9 state, raising the possibility of a nickel analog to the cuprates (9). At the same time, key differences between the systems have been proposed both early on (10) and more recently following the successful demonstration of nickelate superconductivity (11–20). Experimentally, however, direct measurements of the electronic landscape of hole-doped infinite-layer nickelates, particularly with variable band filling, are lacking.In the family of superconducting cuprates, extensive spectroscopic characterization of bulk samples enabled a thorough understanding of their electronic structure, especially regarding the role of doped holes (6). Compared to such bulk phase systems, the successful realization of nickelate superconductivity in epitaxial thin films complicates their characterization by traditional bulk techniques, instead requiring local probes as well as measurements with high sensitivity due to the metastable nature of these compounds. Here, we directly probe the electronic structure across the doped Nd1−xSrxNiOy (y = 2, 3) family using locally resolved electron energy-loss spectroscopy (EELS) in the scanning transmission electron microscope (STEM) highly optimized for detection of subtle spectroscopic signatures. We identify a distinct prepeak feature in the NdNiO3 O-K edge, which disappears completely in NdNiO2, consistent with the predicted Mott–Hubbard character of the infinite-layer parent compound (10, 18). Moreover, our systematic study across a series of superconducting hole-doped Nd1−xSrxNiO2 (x≤ 0.225) films uncovers an emergent feature in the O 2p band with hole doping, reminiscent of the Zhang–Rice singlet (ZRS) peak in the isostructural superconducting cuprates. Unlike the cuprates, however, the spectral weight of the O 2p feature is small even at high doping levels, suggesting necessary involvement of other bands. Indeed, parallel spectroscopy of both Ni and Nd confirms contributions of the Ni 3d and modification of the Nd 5d states, demonstrating, in contrast to the superconducting cuprates, the multiband nature of this system.Infinite-layer nickelate thin films are stabilized by soft-chemistry topotactic reduction (21) of the perovskite phase grown epitaxially on SrTiO3 (001) substrates and capped with a thin layer of SrTiO3 by pulsed-laser deposition (PLD), as previously described (8, 22). Capping layers between 2 and 25 nm have been found to stabilize and support the infinite-layer structure during topotactic reduction, producing more crystallographically uniform films (22). STEM-EELS is therefore an ideal way to spectroscopically probe the nickelate thin film without contributions from the capping layer (or substrate). Starting with the undoped NdNiO2 film, atomic-resolution high-angle annular dark-field (HAADF) STEM imaging confirms the overall high crystalline quality of the film (Fig. 1a) after optimization of the epitaxial growth and subsequent reduction process as discussed elsewhere (22). Elemental EELS mapping performed in the STEM confirms the abruptness of the nickelate–SrTiO3 interface and shows no obvious impurity phases (Fig. 1b). Extended defects are, however, present and must be taken into account when interpreting more commonly used area-averaged spectroscopic measurements (SI Appendix, Figs. S1 and S2). Spatially resolved STEM-EELS reveals distinct O-K near-edge fine structures in different regions of the same film including some crystalline defects and small pockets that have not completely reduced to NdNiO2 (Fig. 1C and SI Appendix, Fig. S2). The intrinsically metastable nature of the infinite-layer compound (21) has required an empirical fine tuning of the conditions for maximal reduction without decomposition (22), so the presence of some unreduced pockets is likely unavoidable at this time. When averaging over the entire film, however, contributions from such variations can mask the electronic character of the pure nickelate phase. In this work, we extract the electronic signatures of the Nd1−xSrxNiOy phases without contributions from extended defects by confining our spectroscopic measurements to crystallographically clean and fully reduced regions of each film using an angstrom-size STEM probe.Open in a separate windowFig. 1.Lattice and electronic structure of NdNiOy thin films. (A) Atomic-resolution HAADF-STEM imaging of an undoped NdNiO2 film shows mostly well-ordered epitaxial structure with some visible crystalline defects. (B) STEM-EELS Nd, Ni, and Ti elemental mapping confirms an abrupt interface with SrTiO3. (C) Observed variations in O-K near-edge fine structure across different regions in the same reduced nickelate film (blue) necessitate local measurements to avoid contributions from defects and partially unreduced regions. For comparison, reference spectra from bulk oxides are also shown. (D) EELS O-K edge of NdNiO2 and a NdNiO3 film before reduction to the infinite-layer phase. The disappearance of the first peak (highlighted in gray), ascribed to metal-oxygen hybridization, indicates filling of the Ni states upon reduction from d7 to d9. (Scale bars, 2 nm.)The distinct characteristic shapes of EELS edges provide access to a wealth of rich chemical information (23–26). To first approximation, the energy loss near-edge structure (ELNES) of core-loss spectra (edges at more than ∼100 eV) reflects the local density of unoccupied states for a specific element and angular momentum. The dipole selection rule permits only transitions from states with orbital angular momentum l to final states with l±1, e.g., s→p or p→d, such that different EELS edges for the same element will probe the unoccupied density of different states. Changes to ELNES shape therefore signify corresponding changes in the density of unoccupied states for a given element that may be affected by hybridization, atomic coordination, or other local bonding effects. Small shifts in the absolute position of the EELS edge can similarly point to shifts or redistribution of unoccupied states. Because EELS probes the energy separation of the unoccupied and the core states, however, shifts in the EELS spectrum can also arise from changes in the core (rather than the unoccupied) level (26, 27). Importantly, the effect of the core-hole created during the excitation process can significantly modify the ELNES and must be taken into account when considering comparisons to ground-state calculations. The interpretation of EELS fine structure provides, in general, similar information as other spectroscopic probes such as unpolarized X-ray absorption spectroscopy (XAS), and can be used to probe the local electronic landscape in great detail.We first explore the electronic character of the parent infinite-layer compound, NdNiO2. In 3d transition metal oxides, the onset structure of the O-K edge is ascribed to hybridization between the O 2p and metal 3d states (28). For NdNiO3, this hybridization results in a strong prepeak feature at ∼527 eV (28–30). Compared to the bulk compound, the prepeak in the NdNiO3 thin film is broadened, consistent with changes to local Ni–O coordination (29, 31) due to epitaxial strain imposed by the SrTiO3 substrate (Fig. 1C). Upon oxygen reduction from the perovskite NdNiO3 to the infinite-layer NdNiO2 phase, the prepeak disappears entirely (Fig. 1D). Here, the suppression of this peak reflects filling of the lower d bands with the change in formal Ni configuration from 3d7 to 3d9 (29) (SI Appendix, Fig. S3). The undoped infinite-layer cuprates exhibit a similar suppression in spectral weight of the d9 states compared to the perovskite phase (32). Notably, the cuprates also show an additional O-K prepeak observed ∼1.5 eV above the d9 peak due to transitions into the d10 upper Hubbard band (32, 33). In the nickelates, however, a similar second peak is not observed, suggesting that it falls within and is thus hidden by the bulk of the O-K edge. Compared to the cuprates, the reduction in nuclear charge from Cu to Ni greatly reduces the covalency of the d9 state and pushes the d10 states to higher energies (12, 14, 18, 34). A d10 peak like that observed in the cuprates would therefore be much weaker and located several electron volts (∼5 to 6 eV) higher in energy, overlapping with the rest of the O-K edge. The lack of a visible spectral peak thus suggests that the d10 states sit at higher relative energies in the nickelates than in the cuprates. This experimental observation is consistent with theoretical predictions that the infinite-layer nickelates have a larger charge-transfer energy Δ than the superconducting cuprates (17, 18).For similar on-site Coulomb interactions U (17, 18), these experiments suggest an electronic structure in the nickelates with a charge-transfer gap Δ larger than U. Using the Zaanen–Sawatzky–Allen classification scheme (35), this places the parent infinite-layer compound in the Mott–Hubbard regime (Δ>U) (11, 17, 18, 36), in contrast to the charge-transfer (Δ<U) cuprates (33) (SI Appendix, Fig. S3). It is, however, likely that hybridization in the nickelates spreads out the distribution of the O 2p bands (37), so that the system actually falls somewhere near the boundary between the two regimes (Mott–Hubbard and charge-transfer) (35). Additionally, while no spectroscopic prepeak is observed in the parent compound NdNiO2, possible effects due to self-doping (17, 34, 38) that may play a role in the detailed electronic structure cannot be directly excluded from the data presented here. Previous results by XAS have similarly suggested the Mott–Hubbard nature of La-based nickelates (36).XAS has also been used to probe the O-K near-edge fine structure of reduced NdNiO2 and unreduced NdNiO3 films on SrTiO3 (34, 36), but these measurements averaged over large parts of the films. The exclusion of defects and the stabilizing SrTiO3 capping layer in our spatially resolved approach explains the apparent discrepancy between these results. In particular, contributions from SrTiO3 or secondary phases in nonoptimized films (22) result in an additional peak at ∼530 eV in the unreduced film as well as a subtle but clear shoulder at the same energy in the reduced film (SI Appendix, Fig. S1). While differences in the perovskite phase are quite striking, below we show that even small contributions from this shoulder in the infinite-layer films can inhibit the ability to probe the effects of hole doping in this system.In addition to the O-K edge, EELS measurements provide simultaneous access to the Ni and Nd states in corresponding localized regions. We measure the Ni-L2,3 edge (Fig. 2A) to probe the Ni 3d states (39) across the same evolution from the perovskite to infinite-layer phase. A clear distinction can be made between the parent NdNiO3 and infinite-layer NdNiO2 thin films, particularly by a broadening and shift of the L3 edge at ∼854 eV (Fig. 2B). The redistribution of spectral weight to higher energies can be seen as a relative decrease in the peak intensity of the L3 peak when spectra are normalized by integration over the full L2,3 edge (Fig. 2A) as well as by the increased width of the L3 peak when spectra are instead normalized by the maximum L3 peak intensity (Fig. 2B). A shift of the infinite-layer L3 peak to higher energies by ∼0.6 eV is also clearly apparent with the latter normalization. In more ionic systems, transition metal L2,3 edges have been used to fingerprint valence states (32, 40–42). In nickel compounds, however, hybridization effects play an important role and are reflected in the fine structure of the Ni-L2,3 edge (27, 39), complicating their physical interpretation. Comparing the two thin-film phases, the asymmetric broadening and long high-energy tails of the Ni-L3 edge in the reduced NdNiO2 suggests increased p-d hybridization (30, 39, 43) consistent with the larger Ni1+ ionic radius and straightening of the Ni–O–Ni bond angles (44) in the infinite-layer structure, as well as with theoretical predictions of strong Ni d orbital hybridization (45). The energy shift of the L3 peak could be reflective of the change in Ni 3d states in the infinite-layer compound, as has been described in other nickel-based systems, but EELS measurements alone cannot rule out the contribution of shifts of the core state (27). More direct interpretation of these spectral changes will require comparison with careful and thorough calculations. Similar to the O-K edge, the spectral differences between bulk and thin film NdNiO3 in the L2,3 are likely due to strain effects imparted by the substrate (46). Unlike the Ni-L2,3 and O-K, the Nd-M4,5 edge (4f states) remains unchanged upon reduction from perovskite to infinite-layer (Fig. 2C), indicating a formal Nd valence of 3+ in both phases (SI Appendix, Fig. S4). This lack of change is consistent with the experimental demonstration of superconducting Pr-based infinite-layer nickelate films, suggesting that the specific configuration of the rare-earth 4f state may be relatively unimportant for superconductivity (47).Open in a separate windowFig. 2.Electronic evolution from perovskite NdNiO3 (black) to infinite-layer NdNiO2 (blue) phase for epitaxial thin films on SrTiO3. (A) The Ni-L2,3 edge (3d states) shows a clear change upon O reduction from NdNiO3 to NdNiO2, including significant broadening and shift of the L3 edge (B). (C) The Nd-M4,5 edge (4f states) shows little or no change between the two compounds. Spectra in A and C are normalized by integrated signal over the full energy ranges shown; the Ni-L3 edges in B have been renormalized by maximum intensity for easier comparison.Having established the nature of the parent compounds, we systematically study the effect of hole doping in a series of superconducting infinite-layer Nd1−xSrxNiO2 films. HAADF-STEM imaging and elemental EELS mapping (Fig. 3 A–C) confirm overall film quality across all samples with uniform distribution of Nd and Ni throughout the films. Summed spectra (Fig. 3D) from clean regions within the films show Sr-M3 signals consistent with the relative nominal Sr doping expected in each case. Transport measurements for the infinite-layer films studied here (Fig. 3 E and ​andF)F) show metallic temperature dependence of the low (x = 0.0, 0.1) Sr (hole)-doped films down to about 50 to 70 K, below which a resistive upturn is observed. Upon increased hole doping (x = 0.2, 0.225), the infinite-layer films become superconducting with transition temperatures Tc, 90%R (the temperatures at which the resistivity is 90% that at 20 K) of 8 and 11 K, respectively (Fig. 3F). The relative values of Tc for these films are consistent with a recent investigation of the doping dependence of Nd1−xSrxNiO2 establishing a superconducting dome spanning 0.125 <x< 0.25 (48).Open in a separate windowFig. 3.Structure and transport of Nd1−xSrxNiO2 films with hole doping x≤ 0.225. (A–C) Atomic-resolution HAADF-STEM and elemental EELS maps of x = 0.10, 0.20, and 0.225 samples, respectively. (D) Sr-M3 EEL spectra from each film show a qualitative trend consistent with the expected nominal Sr doping in each sample. A similar spectrum from SrTiO3 is shown in gray for comparison. (E and F) Resistivity vs. temperature curves for the four Nd1−xSrxNiO2 films studied here. Lightly Sr-doped (x = 0.0, 0.1) films show metallic behavior with a resistive upturn near 50 to 70 K, while more strongly doped (x = 0.2, 0.225) films show superconducting transitions (Tc,90%R) at 8 and 11 K, respectively. Data for x = 0.0 are from ref. 8.In the O-K edge, we observe an emergent feature at ∼528 eV as a function of Sr (hole) doping across the Nd1−xSrxNiO2 films (Fig. 4A). The strength of this peak increases with hole doping (Fig. 4B), x, although the spectral weight of the x = 0.2 sample is higher than the more strongly doped x = 0.225 sample, likely due to the inclusion of some Ruddlesden–Popper fault regions (SI Appendix, Figs. S1 and S5). A similar spectral feature is also observed upon hole doping in the cuprates (33, 49). As in the cuprates, we attribute the emergent spectral feature to doped holes in the oxygen orbitals forming 3d9L_ states, where L_ is the oxygen “ligand” hole (4, 49). The coordination of two such holes into a ZRS state is thought to reduce the cuprate system into an effective single-band model (4, 50). In nickelates the picture is less clear: despite the similarity in presence and doping dependence of this O-K edge feature in both systems, a few key spectroscopic differences are apparent. In the cuprates, hole doping results in a tradeoff of spectral weight from the d10 upper Hubbard band into the emergent d9L_ states (33). In the nickelates, we do not observe a similar trend, most likely due to shift of the d10 band to much higher energy, as described by the Mott–Hubbard picture discussed above. Perhaps more importantly, the overall strength of the hole doping (3d9L_) feature is significantly reduced in the nickelates compared to the cuprates, even for comparatively high levels of hole doping. The weak effect of doping on the O-K edge raises the question: where else do the holes go?Open in a separate windowFig. 4.Electronic structure evolution of Nd1−xSrxNiO2 with hole doping x≤ 0.225. (A) A spectral feature at ∼528 eV emerges with increased Sr (hole) doping in the O-K edge, attributed to d9L_ states. The dotted line marks the d9 peak position in NdNiO3. (B) The integrated signal over the d9L_ feature spectral range (527 to 529 eV), highlighted gray in A, increases with hole doping. (C and D) The Ni-L3 edge systematically shifts to high energies and broadens with hole doping. The dotted line marks the L3 peak position in NdNiO3. (E) The Nd-N2,3 edge (5d states) shows a small shift to higher energies upon Sr doping of the perovskite phase. A further shift is observed upon O reduction to infinite-layer Nd0.8Sr0.2NiO2. The dotted line marks the N2,3 peak position in NdNiO3. For comparison, reference spectra are plotted in gray for NdNiO2 in A and C and for NdNiO3 in E.The ability of EELS measurements to access several edges simultaneously allows us to consider the corresponding effects on the Ni-L2,3 edge (3d bands) (Fig. 4C). The Ni-L3 peak position (Fig. 4D and SI Appendix, Fig. S6) shifts systematically to higher energies upon hole doping, with a total increase of ∼0.3 eV between the undoped x = 0.0 and the x = 0.225 films. We also observe the L3 edge broadening as an extended tail in the doped samples (SI Appendix, Fig. S5). When compared to previously reported spectra for various Ni formal valences (30, 39, 43, 44, 51), the changes in the Ni-L2,3 edge may suggest increased hybridization and a mixed-valent Ni state. Similar to the evolution from perovskite to infinite layer, we interpret the doping-dependent broadening of the L3 edge as a sign of increasing hybridization of the Ni 3d states with the introduction of more holes through Sr doping.We also probe the Nd-N2,3 edge (5d states) across the sample series (Fig. 4E). Given the extreme technical challenges due to the need for simultaneous high-spatial resolution, high-energy resolution, and high signal-to-noise ratio (SNR), current data are limited to only the compositions shown in Fig. 4E, although characterizing the full doping dependence of the N2,3 edge is the subject of ongoing work. The x = 0.2 hole-doped perovskite shows a ∼200-meV shift to higher energies compared to the parent compound. In the x = 0.2 infinite-layer film, we observe a twofold increase in this energy shift as well as an increase in spectral intensity. We note that a number of complicating factors could also contribute to changes in the edge shape, such as hybridization with the O states and a modified Madelung potential, both of which may be dominated simply by the stark difference in the crystalline environment of Nd from the perovskite to infinite-layer phase. Calculations by Choi et al. (13) have also suggested that Nd 5d bands may be broadened by disorder in the 4f states in the hole-doped infinite-layer compound. Others have predicted that these small 5d contributions can and, in fact, should be avoided by adjusting the choice of rare earth to shrink the corresponding Fermi pockets (11, 14). Given the ongoing theoretical debate about the importance of the parent cation 5d bands (10, 19, 52), direct experimental measurements of these states provide valuable information, although full interpretation of the observed shifts in the Nd-N2,3 edge will require detailed calculations.Together, the evolution of all three edges (O-K, Ni-L2,3, and Nd-N2,3) suggest markedly different behavior of doped holes in infinite-layer nickelates than in the cuprates. Rather than doping mostly onto O, our data are consistent with a picture in which the large charge-transfer energy Δ>U of the Mott–Hubbard regime instead pushes holes also onto Ni, resulting in d8-like states above the O 2p band. Hybridization between the ligand holes and d8 states results in the observed d9L_ peak at ∼528 eV in the O-K edge. Hybridization with this d9L_ pushes the d8 state up toward the Fermi energy, resulting in a “mixed-valent metal” state, as also suggested by the observed broadening of the Ni-L2,3 edge. Finally, the effects measured in the Nd-N2,3 edge illustrate some interaction with the doped holes in the Nd 5d bands as well, consistent with a multiband system fundamentally different from the effective single-band superconducting cuprates, as previously proposed by theoretical models (12, 16, 20, 45) and Hall effect measurements (48).Beyond hole doping, the large parameter space, including rare-earth cation (11, 14), epitaxial strain, reduction process, etc., yet to be explored for this new family of superconductors will provide a rich backdrop for future experiments. Especially in thin-film form, spatially resolved spectroscopy will be a powerful technique to not only minimize spectral signatures due to defects but also exclude contributions from stabilizing capping layers (22).The data presented here are direct evidence of key differences between the electronic structures of two isostructural superconducting oxide families. Analysis of the parent perovskite and infinite-layer compounds establishes the Mott–Hubbard character of NdNiO2, distinct from the charge-transfer cuprates. Within the doped infinite-layer series, we observe effects of hole doping in not just the O 2p but also the Ni 3d and Nd 5d bands, drawing a further distinction between the nickelates and their isostructural cuprate cousins. Lacking both the charge-transfer character and magnetic order (53) of the cuprates, understanding the nature of superconductivity in this new family of nickelates will require new insights from both theory and experiment.